Journal of Modern Physics, 2019, 10, *-* http://www.scirp.org/journal/jmp ISSN Online: 2153-120X ISSN Print: 2153-1196 DOI: 10.4236/***.2019.***** **** **, 2019 1 Journal of Modern Physics The Solution Cosmological Constant Problem Jaykov Foukzon1, Elena Men'kova2, Alexander Potapov3 1Department of Mathematics, Israel Institute of Technology, Haifa, Israel 2All-Russian Research Institute for Optical and Physical Measurement, Moscow, Russia 3Kotel'nikov Institute of Radioengineering and Electronics, Russian Academy of Sciences, Moscow, Russia Email: jaykovfoukzon@list.ru, e_menkova@mail.ru, potapov@cplire.ru Abstract The cosmological constant problem arises because the magnitude of vacuum energy density predicted by the Quantum Field Theory is about 120 orders of magnitude larger then the value implied by cosmological observations of accelerating cosmic expansion. We pointed out that the fractal nature of the quantum space-time with negative Hausdorff-Colombeau dimensions can resolve this tension. The canonical Quantum Field Theory is widely believed to break down at some fundamental high-energy cutoff ∗Λ and therefore the quantum fluctuations in the vacuum can be treated classically seriously only up to this high-energy cutoff. In this paper we argue that the Quantum Field Theory in fractal space-time with negative Hausdorff-Colombeau dimensions gives high-energy cutoff on natural way. We argue that there exists hidden physical mechanism which cancels divergences in canonical 4 4,QED QCD , Higher-Derivative-Quantum gravity, etc. In fact we argue that corresponding supermassive Pauli-Villars ghost fields really exist. It means that there exists the ghost-driven acceleration of the universe hidden in cosmological constant. In order to obtain the desired physical result we apply the canonical Pauli-Villars regularization up to ∗Λ . This would fit in the observed value of the dark energy needed to explain the accelerated expansion of the universe if we choose highly symmetric masses distribution between standard matter and ghost matter below the scale ∗Λ , i.e., ( ) ( ). . eff eff, , ,s m g mf f mc cμ μ μ μ μ μ ∗− = ≤ < Λ The small value of the cosmological constant is explained by tiny violation of the symmetry between standard matter and ghost matter. Dark matter nature is also explained using a common origin of the dark energy and dark matter phenomena. Keywords Cosmological Constant Problem, Quantum Field Theory, Vacuum Energy Density, Quantum Space-Time, Hausdorff-Colombeau Dimension, Quantum How to cite this paper: Author 1, Author 2 and Author 3 (2019) Paper Title. Journal of Modern Physics, 10, *-*. https://doi.org/10.4236/***.2019.***** Received: **** **, *** Accepted: **** **, *** Published: **** **, *** Copyright © 2019 by author(s) and Scientific Research Publishing Inc. This work is licensed under the Creative Commons Attribution International License (CC BY 4.0). http://creativecommons.org/licenses/by/4.0/ Open Access J. Foukzon et al. DOI: 10.4236/***.2019.***** 2 Journal of Modern Physics Fluctuations, High-Energy Cutoff, Canonical Pauli-Villars Regularization, Universe 1. Introduction 1.1. The Formulation of the Cosmoloigical Constant Problem The cosmological constant problem arises at the intersection between general relativity and quantum field theory, and is regarded as a fundamental unsolved problem in modern physics. A peculiar and truly quantum mechanical feature of the quantum fields is reminded that they exhibit zero-point fluctuations everywhere in space, even in regions which are otherwise "empty" (i.e. devoid of matter and radiation). This vacuum energy density is believed to act as a contribution to the cosmological constant λ appearing in Einstein's field equations from 1917, 4 1 8π 2 GR g R T cμν μν μν ′− = (1) where Rμν and R refer to the curvature of space-time, gμν is the metric, Tμν′ is the the energy-momentum tensor, 4 1 0 0 0 0 1 0 0 0 0 1 08π 0 0 0 1 cT T Gμν μν λ    − ′ = +  −     (2) where Tμν is the energy-momentum tensor of matter. Thus 00 00T T λε′ = + , T T Pαβ αβ αβ λδ′ = + , where 4 8π .P c Gλ λε λ= − = (3) Reminding that under Lorentz transformations ( ) ( ), , ,P P Pλ λ λ λ λ λε ε ε′ ′→ → the quantities λε and Pλ are changes by the law 2 2 2 2, .1 1 P P Pλ λ λ λλ λ ε β β ε ε β β + +′ ′= = − − (4) Thus for the quantities λε and Pλ Lorentz invariance holds by Equation (3) [1]. In modern cosmology it is assumed that the observable universe was initially vacuumlike, i.e., the cosmological medium was non-singular and Lorentz invariant. In the earlier, non-singular Friedmann cosmology, the Friedmann universe comes into being during the phase transition of an initial vacuumlike state to the state of "ordinary" matter [2] [3]. The Friedmann equations start with the simplifying assumption that the universe is spatially homogeneous and isotropic, i.e. the cosmological principle; empirically, this is justified on scales larger than 126, ~100 Mpc. The cosmological principle implies that the metric of the universe must be of the form RoJ. Foukzon et al. DOI: 10.4236/***.2019.***** 3 Journal of Modern Physics bertson-Walker metric [2]. Robertson-Walker metric reads ( ) ( ) 2 2 2 2 2 2 2 2 2 dd d d sin d . 1 rs t a t r kr θ θ φ   = − + + −  (5) For such a metric, the Ricci curvature scalar is 6R k= − and it is said that space has the curvature k. The scaling factor ( )a t rescales this curvature for a given time t, producing a curvature ( ) ( )k t k a t= . The scaling factor ( )a t is given by two independent Friedmann equations for modeling a homogeneous, isotropic universe reads ( )2 2 , 3 3 6 G Ga a k a pε ε= − = − +  (6) and the equation of state ( ) ,p p ε= (7) where p is pressure and ε is a density of the cosmological medium. For the case of the vacuumlike cosmological medium equation of state reads [2] [3] [4]: .p ε= − (8) By virtue of Friedman's Equations (1.1.6) in the Universe filled with a vacuum-like medium, the density of the medium is preserved, i.e. constε = , but the scale factor ( )a t grows exponentially. By virtue of continuity, it can be assumed that the admixture of a substance does not change the nature of the growth of the latter, and the density of the medium hardly changes. This growth, interpreted by analogy with the Friedmann models as an expansion of the universe, but almost without changing the density of the medium! was named inflation. The idea of inflation is the basis of inflation scenarios [2]. Non-singular cosmology [2] [4] suggests that the initial state of the observable universe was vacuum-like, but unstable with respect to the phase transition to the ordinary non-Lorentz-invariant medium. This, for example, takes place if, by virtue of the equations of state of the medium, a fluctuation decrease in its density d violates the condition of vacuum-like degeneration, p ε= − or, which is the same, 3 2 0p ε ε+ = − < , replacing it with 2 3 0.pε ε− < + < (9) According to Friedman's equations, it corresponds to an accelerated expansion of the cosmological medium, accompanied by a drop in its density, which makes the process irreversible [2]. The impulse for expansion in this scenario, the vacuum-like environment, is not reported to itself (bloating), but to the emerging Friedmann environment. In review [5], Weinberg indicates that the first published discussion of the contribution of quantum fluctuations to the cosmological constant was a 1967 paper by Zel'dovich [6]. In his article [1] Zel'dovich emphasizes that zeropoint energies of particle physics theories cannot be ignored when gravitation is taken into account, and since he explicitly discusses the discrepancy between estimates of vacuum energy and observations, he is clearly pointing to a cosmological conJ. Foukzon et al. DOI: 10.4236/***.2019.***** 4 Journal of Modern Physics stant problem. As well known zeropoint energy density of scalar quantum field, etc. is divergent ( ) ( ) 2 2 2 2 vac 3 0 2π d . 2π cm p m c p pε ∞ = + = ∞∫  (10) In order to avoid difficulties mentioned above, in article [1] Zel'dovich has applied canonical Pauli-Villars regularization [7] [8] and formally has obtained a finite result (his formulas [1], Eqs. (VIII.12)-(VIII.13) p. 228) ( ) ( ) 4 4 vac vac 0 1 ln d , 8 8π cp f G λε μ μ μ μ ∞ = − = =∫ (11) where ( ) ( ) ( )2 4 0 0 0 d d d 0.f f fμ μ μ μ μ μ μ μ ∞ ∞ ∞ = = =∫ ∫ ∫ (12) Remark 1.1.1. Unfortunately, Equation (11) and Equation (12) give nothing in order to obtain desired numerical values of the zero-point energy density ε . In his paper [1], Zel'dovich arrives at a zero-point energy (his formula (IX.1)) 3 17 3 10 2 vac ~ 10 g cm , ~ 10 cm , mcmε λ − − =     (13) where m (the ultra-violet cut-of) is taken equal to the proton mass. Zel'dovich notes that since this estimate exceeds observational bounds by 46 orders of magnitude it is clear that "... such an estimate has nothing in common with reality". In his paper [1], Zel'dovich wrote: Recently A. D. Sakharov proposed a theory of gravitation, or, more precisely, a justification GR equation based on consideration of vacuum fluctuations. In this theory, the essential assumption is that there is some elementary length L or the corresponding limiting momentum 0p L=  . Shorter lengths or for large impulses theory is not applicable. Sakharov gets the expression of gravitational constant G through L or 0p (his formula (IX.6)) 3 2 3 2 0 .c L cG p = =   (14) This expression has been known since the days of Planck, but it was read "from right to left": gravity determines the length L and the momentum 0p . According to Sakharov, L and 0p are primary. Substitute (IX. 6) in the expression (IX. 4), we get 6 5 6 7 vac vac2 3 2 3 0 0 , .m c m c p p ρ ε= =   (15) That is expressions that the first members (in the formulas (VIII.10), (VIII. 11)) which are vanishes (with 0p →∞ ). Thus, we can suggest the following interpretation of the cosmological constant: there is a theory of elementary particles, which would give (according to the mechanism that has not been revealed at the present time) identically zero vacuum energy, if this theory is applicable infinitely, up to arbitrarily large momentum; there is a momentum 0p , beyond J. Foukzon et al. DOI: 10.4236/***.2019.***** 5 Journal of Modern Physics which the theory is non applicable; along with other implications, modifying the theory gives different from zero vacuum energy; general considerations make it likely that the effect is portional 20p − . Clarification of the question of the existence and magnitude of the cosmological constant will also be of fundamental importance for the theory of elementary particles. Nonclassical Assumptions (I) In contrast with Zel'dovich paper [1] we assume that Poincaré group is deformed at some fundamental high-energy cutoff ∗Λ [9] [10] [11] in accordance with the basis of the following deformed Poisson brackets { } ( ) { } { } 1 0 0 1 0 , , , 0, , x x x x p p x p p μ ν μ ν ν μ μ ν μ ν μν μ ν η η η η − − = − = = − +   (16) where , 0,1, 2,3μ ν = , ( )1, 1, 1, 1μνη = + − − − and is a parameter identified as the ratio between the high-energy cutoff ∗Λ and the light speed. The corresponding to (16) momentum transformation reads [11] ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 2 00 0 1 1 0 0 1 1 0 0 , , 1 1 1 1 , , 1 1 1 1 xx x x x y z y z x x p up cp up p p c p up c p up p pp p c p up c p up γγ γ γ γ γ γ γ γ γ − − − − −− ′ ′= = + − − + − −       ′ ′= = + − − + − −           (17) and coordinate transformation reads [11] ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 2 1 1 0 0 1 1 0 0 , , 1 1 1 1 , , 1 1 1 1 x x x x t ux c x ut t x c p up c p up y zy z c p up c p up γ γ γ γ γ γ γ γ γ γ − − − − − − ′ ′= = + − − + − −       ′ ′= = + − − + − −           (18) where 2 21 u cγ = − . It is easy to check that the energy E c=  , identified as the high-energy cutoff ∗Λ , is an invariant as it is also the case for the fundamental length l c E ∗Λ = =   . Remark 1.1.2. Note that the transformation (17) defined in p-space and the transformation (1.1.18) defined in x-space becomes Lorentz for small energies and momenta and defines a large invariant energy 1l ∗ − Λ . The high-energy cutoff ∗Λ is preserved by the modified action of the Lorentz group [9] [10]. This meant that the canonical concept of metric as quadratic invariant collapses at high energies, being replaced by the non-quadratic invariant [9]: ( ) 2 0 , 1 ab a bp pp l p η Λ∗ = + (19) or by the non-quadratic invariant ( ) 2 0 , 1 ab a bp pp l p η Λ∗ = − (20) where 1, , 0,1, 2,3l a b ∗ − Λ ∗= Λ = . J. Foukzon et al. DOI: 10.4236/***.2019.***** 6 Journal of Modern Physics Remark 1.1.3. Note that: 1) the invariant (16) is infinite for the new negative invariant energy scale of the theory 1l ∗ − ∗ ΛΛ = − , and it's not quadratic for energies close or above and 2) the invariant (17) is infinite for the new positive invariant energy scale of the theory 1l ∗ − ∗ ΛΛ = . Remark 1.1.4. It is also clear from Equation (16) and Equation (17) that the symmetry of positive and negative values of the energy is broken. The two theories with the two signs of lΛ obviously are physically distinct; and we know of no theoretical argument which fixes the sign of lΛ The massive particles have a positive invariant 2 0p > which can be identified with the square of the mass 2 2p m= , ( 1c = ). Thus in the case of the invariant (16) we obtain ( ) ( ) 2 2 2 10 02 0 , , 1 p p m p l l p ∗ ∗ − Λ Λ − = ∈ − ∞ + (21) From Equation (18) we obtain ( ) 2 4 2 2 2 0 2 2 2 22 2 1 . 1 11 m l m l p p m m l m lm l ∗ ∗ ∗ ∗∗ Λ Λ Λ ΛΛ = + + + − −− (22) In the case of the invariant (17) we obtain ( ) ( ) 2 2 2 10 02 0 , , . 1 p p m p l l p ∗ ∗ − Λ Λ − = ∈ −∞ − (23) From Equation (20) we obtain ( ) 2 4 2 2 2 0 2 2 2 22 2 1 1 11 m l m l p p m m l m lm l ∗ ∗ ∗ ∗∗ Λ Λ Λ ΛΛ = − − + + − −− (24) The action for a scalar field φ must be invariant under the deformed Lorentz transformations. The invariant action reads [10] ( )( ) 24 2 0 1 d . 2 21 ab a b mS x l η φ φ φ φ ∗Λ ∂ ∂ = +  + ∂  ∫ (25) Thus there is no linear field equation. Remark 1.1.5.Throughout this paper, we use below high-energy cutoff ∗Λ the perturbative expansion ( )( ) ( ) 2 4 21 d . 2 2 ab a b mS x O lη φ φ φ ∗Λ   = ∂ ∂ + +    ∫ (26) and dealing in Lorentz invariant approximation ( )( ) 2 4 21 d . 2 2 ab a b mS x η φ φ φ   ∂ ∂ +    ∫ (27) since for 1l ∗Λ  the expansion (26) holds. J. Foukzon et al. DOI: 10.4236/***.2019.***** 7 Journal of Modern Physics (II) The canonical concept of Minkowski space-time collapses at a small distance 1l ∗ − Λ ∗= Λ to fractal space-time with Hausdorff-Colombeau negative dimension and therefore the canonical Lebesgue measure 4d x being replaced by the Colombeau-Stieltjes measure with negative Hausdorff-Colombeau dimension D− : ( )( ) ( )( )( )4d , d ,x v s x xεε εη ε = (28) where ( )( )( ) ( ) 1 D v s x s xε ε ε ε − −  = +      and ( )s x x xμμ= , see Section 3 and [12]. (III) The canonical concept of momentum space collapses at fundamental high-energy cutoff ∗Λ to fractal momentum space with Hausdorff-Colombeau negative dimension and therefore the canonical Lebesgue measure 3d k , where ( ), ,x y zk k k=k being replaced by the Hausdorff-Colombeau measure ( ) ( ) ( ) 1, d dd , D D D D D D D D D p p p ε ε ε ε + + + − − − − + − −∆ ∆ ∆ =    + +        k k k (29) where ( ) ( )22π 2DD D±± ±∆ = Γ and x y zp k k k= = + +k and where 6D D+ −− ≤ − , see Section 3 and ref. [9]. Hausdorff-Colombeau measure (29) avoids classical divergence (10) of the zeropoint energy ( )vac mε and instead Equation (10) one obtains ( ) ( ) ( ) 2 2 3 2 2 2 4 vac 0 , d d . p p D p m p m p p m D D pp p p ε ε ε ∗ −∗ ∞+ − ∗ ∗ + = + + ∆ ∆  +    ∫ ∫  (30) See Section 5 and ref. [12]. Remark 1.1.6. If we take the Planck scale (i.e. the Planck mass) as a cut-off, the vacuum energy density ( )vac ,p mε ∗ is 10121 times larger than the observed dark energy density deε . Several possible approaches to the problem of vacuum energy have been discussed in the contemporary literature, for the review see [5], [12]. They can be roughly devided into four different groups: 1) Modification of gravity on large scales. 2) Anthropic principle. 3) Symmetry leading to vac 0ε = . 4) Adjustment mechanism, see. 5) Hidden nonstandard dark matter sector and corresponding hidden symmetry leading to vac 0ε  , see [12]. (IV) We assume that there exists the nonstandard dark matter sector formed by ghost particles, see [12]. 1.2. Zel'dovich Approach by Using Pauli-Villars Regularization Revisited. Ghosts as Physical Dark Matter Remind that vacuum energy density for free scalar quantum field is J. Foukzon et al. DOI: 10.4236/***.2019.***** 8 Journal of Modern Physics ( ) ( ) ( )2 2 2 2 2 23 0 0 1 4π d d , 2 2π c p p p K p p p KIε μ μ μ μ ∞ ∞ = + = + =∫ ∫  (31) where 0m cμ = . From Equation (31) one obtains [1] ( ) ( ) 4 0 2 2 d . 3 K p pp KF p μ μ μ ∞ = = + ∫ (32) For fermionic quantum field one obtains ( ) ( ) ( ) ( ), 4 .KI p KFε μ μ μ μ= = − (33) Thus free vacuum energy density ε and corresponding pressure p is ( ) ( ), .i i i i i i C I P C Fε μ μ= =∑ ∑ (34) From Equation (34) by using Pauli-Willars regularization [7] [8] in general case one obtains [1] ( ) ( ) ( ) ( )d , d .f I P f Fε μ μ μ μ μ μ= =∫ ∫ (35) In order to obtain asymptotical expansion on the parameter 0p of the quantity ( ) 0 3 2 2vac 0 0, d p p m p p mε = +∫ let us evaluate now the following integral ( ) 0 0 0 2 2 2 2 2 2 2 2 2 0 0 0 2 2 2 2 2 3 3 2 2 0 , d d d d 1 d 1 d pp p p p p p p p I p p p p p p p p p p p p p p p p p p p μ μ μ μ μ μ μ μ μ μ μ μμ = + = + + + = + = + + + ∫ ∫ ∫ ∫ ∫ ∫ (36) and ( ) 0 0 0 4 4 4 0 2 2 2 2 2 2 0 0 4 3 2 2 2 0 2 1 d 1 d 1 d, 3 3 3 1 d 1 d , 3 3 1 pp p p p p p p p p p p pF p p p p p p p p p p μ μ μ μ μ μ μ μ μ μ = = + + + + = + + + ∫ ∫ ∫ ∫ ∫ (37) where , 1, 1 1p r r p rμ μ μ= > < < . Note that 2 2 4 6 2 2 4 6 2 4 6 2 2 2 3 3 2 2 3 1 1 11 1 2 8 16 1 1 11 2 8 16 p p p p p p p p p pp p μ μ μ μ μ μ μμ μ + = + − + + + = + = + − + +   (38) By inserting Equation (38) into Equations (36) one obtains ( ) ( ) 6 4 4 2 2 4 5 80 0 1 0 0 02 0 1 1 1 1, ln , 4 4 8 32 p I p C p p p O p μμ μ μ μ μ μ − = + + − − +    (39) where 4 2 2 21 0 d p C p p pμμ μ= +∫ . Note that J. Foukzon et al. DOI: 10.4236/***.2019.***** 9 Journal of Modern Physics 1 2 2 4 6 2 2 4 6 1 3 51 1 2 8 16p p p p μ μ μ μ −    + = − + − +      (40) By inserting Equation (40) into Equation (37) one obtains ( ) ( ) 6 4 4 2 2 4 5 80 0 2 0 0 02 0 1 1 1 5, ln . 12 12 8 32 p F p C p p p O p μμ μ μ μ μ μ − = + − + + +    (41) By inserting Equation (39) and Equation (41) into Equations (35) one obtains ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff eff eff eff eff 4 2 2 4 0 0 1 0 0 0 0 4 6 8 5 02 0 0 00 4 2 2 4 0 0 2 0 0 0 0 1 1 1d d ln d 4 4 8 1 1 1ln d d , 8 32 1 1 1d ln d 12 12 8 1 8 p f p f C p f f f O f p p p p f p f d C p f μ μ μ μ μ μ μ μ μ ε μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ −  = + + −       + − +          = − + +    − ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ( ) ( ) ( ) ( ) eff eff eff 4 6 8 5 02 0 0 00 5 1ln d d . 32 f f O f p p μ μ μ μ μ μ μ μ μ μ μ μ −    + +         ∫ ∫ ∫ (42) We choose now ( ) ( ) ( ) eff eff eff 2 4 0 0 0 d d d 0.f f f μ μ μ μ μ μ μ μ μ μ μ= = =∫ ∫ ∫ (43) By inserting Equation (43) into. Equations (42) one obtains ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff 4 2 eff 0 0 4 2 eff 0 0 1 ln d , 8 1 ln d . 8 f O p p f O p μ μ ε μ μ μ μ μ μ μ μ μ μ − − = + = − + ∫ ∫ (44) Taking the limit p →∞ in Equation (44) gives ( ) ( ) ( ) ( ) ( ) ( ) eff eff 4 eff 0 4 eff 0 1 ln d , 8 1 ln d . 8 f p f μ μ ε μ μ μ μ μ μ μ μ μ μ = = − ∫ ∫ (45) Thus finally we obtain [3] ( ) ( ) ( ) ( ) eff 4 4 eff eff 0 1 ln d . 8 8π cp f G μ ε μ μ μ μ μ μ Λ= − = =∫ (46) Remark 1.2.1. Remind that Pauli-Villars regularization consists of introducing a fictitious mass term. For example, we would replace a propagator ( )2 201 k m i− +  , by the regulated propagator ( )2 2 2 2 2 2 20 0 0 1 ,N Ni ii i i i a a k k m i k m i k m i= = ∆ = = − − + − + − +∑ ∑   (47) where 0 1a = and , 1, 2, ,im i N=  can be thought of as the mass of a fictitious heavy particle, whose contribution is subtracted from that of an ordinary particle. J. Foukzon et al. DOI: 10.4236/***.2019.***** 10 Journal of Modern Physics Assume that 2 2 1im k < , if we expand each term of this sum (46) as a power series in 2k i+  we get ( ) ( ) ( ) 2 2 2 2 30 0 02 2 1 .N N Ni i ii i i a a m k O k i k i k i = = =    ∆ = + +  +  + +  ∑ ∑ ∑   (48) For a renormalizable theory the maximum supercriticial power of divergence of any integral is quadratic, so that the ( )61O k terms are ultraviolet finite. The finiteness of the regulated integral is then guaranteed by requiring that 2 0 00, 0. N N i i ii ia a m= == =∑ ∑ (49) Remark 1.2.2. Note that in order to apply Pauli-Villars regularization to QFT with Lagrangian ( ), , ,μ μφ ψ φ ψ∂ ∂L we would replace the Lagrangian ( ), , ,μ μφ ψ φ ψ∂ ∂L by Lagrangian ( ), , ,μ μφ ψ φ ψ∂ ∂L , where [7]: ( ) ( ) ( ) ( ) ( ) ( )2 2, , , ,n n n n n nn nx x b x x x c xφ φ φ μ ψ ψ ψ= + = +∑ ∑   (50) where commutator for nφ  and anticommutator for nψ  reads ( ) ( ) ( ) ( ) ( ){ } ( ) 2 2 2 2 2 2 , , , , , , , , , . m m n n n n mn m m n n n n mn x x i x x x x i S x x φ μ φ μ ρ μ δ ψ ψ ε δ  ′ ′= − ∆ −  ′ ′= − −       From Equations (50)-Equations (51) one obtains ( ) ( ) ( ) ( ) ( ) ( ) 2 2 0 2 0 , , , , , . N n n nn N n n n nn x x i b x x x x i c c S x x φ φ ρ μ ψ ψ ε = =  ′  ′= ∆ −   ′  ′= − −  ∑ ∑  (52) Assume now that 2 2 2 2 0 0 0 00, 0, 0, 0. N N N N n n n n n n n n n n n nn n n nb b c c c cρ ρ μ ε ε= = = == = = =∑ ∑ ∑ ∑  (53) From Equations (53) it follows directly that QFT with Lagrangian ( ), , ,μ μφ ψ φ ψ∂ ∂L is finite QFT with indefinite metric [4], see Remark 1.2.1. Remark 1.2.3. Note that "bad ghosts" represent general meaning of the word "ghost" in theoretical physics: states of negative norm [7] or fields with the wrong sign of the kinetic term, such as Pauli--Villars ghosts φ , whose existence allows the probabilities to be negative thus violating unitarity. The quadratic lagrangian 2φL for φ begins with a wrong sign kinetic term [in ( + − − − ) signature] 2 1 2 μ φ μφ φ= − ∂ ∂ +L (54) Remark 1.2.4. Note that in order to obtain Equations (44), the standard quantum fields do not need to couple directly to the ghost sector. In this paper the ghost sector is considered as physical mechanism which acts only on a function ( )f μ in Equations (43). It means that there exists the ghost-driven acceleration of the universe hidden in cosmological constant Λ . Remark 1.2.5. As pointed out in paper [13] even if the standard model fields have no direct couplings to the ghost sector, they will indirectly interact with it J. Foukzon et al. DOI: 10.4236/***.2019.***** 11 Journal of Modern Physics through gravity, and the propagation of gravity through the ghost condensate gives rise to a fascinating modification of gravity in the IR. However, no modifications of gravity can be seen directly, and no cosmological experiment can distinguish the ghost-driven acceleration from a cosmological constant. Remark 1.2.6. In order to obtain desired physical result from Equations (45), i.e., 29 3 47 4 3 5 vac 0.7 10 g cm 2.8 10 GeV cε − − −= × ⋅ = ×  (55) we assume that ( ) ( ) ( ). . . . ,s m g mf f fμ μ μ= + (56) where ( ). .s mf μ corresponds to standard matter and where ( ). .g mf μ corresponds to a physical ghost matter. Remark 1.2.7. We assume now that ( ) ( ) eff eff , 1 0 nO n f μ μ μ μ μ μ − > ≤=  > (57) From Equation (57) and Equation (45) it follows directly that ( ) ( ) ( ) ( ) ( ) eff 4 5 eff eff eff eff 0 1 ln d ln . 8 np f O μ μ ε μ μ μ μ μ μ μ− += = ≤∫ (58) Remark 1.2.8. However serious problem arises from non-renormalizability of canonical quantum gravity with Einstein-Hilbert action 41 d . 16πEH S x gR G = −∫ (59) For example taking 3Λ particles of energy a per unit volume gives the gravitational self-energy density of order 6Λ , i.e., the density εΛ diverges as 6Λ 6 ,GεΛ Λ (60) where Λ is a high-energy cutoff [5]. In order to avoid these difficulties we apply instead Einstein-Hilbert action (59) the gravitational action which includes terms quadratic in the curvature tensor ( )4 2 2d 2 ,x g R R R Rμνμνα β κ −I = − − − +∫ (61) Remark 1.2.9. Gravitational actions (61) which include terms quadratic in the curvature tensor are renormalizable [14]. The requirement that the graviton propagator behaves like 4p− for large momenta makes it necessary to choose the indefinite-metric vector space over the negative-energy states. These negative-norm states cannot be excluded from the physical sector of the vector space without destroying the unitarity of the S matrix, however, for their unphysical behavior may be restricted to arbitrarily large energy scales ∗Λ by an appropriate limitation on the renormalized masses 2m and 0m . Remark 1.2.10. We assume that 0 eff 2 eff,m c m cμ μ  . Remark 1.2.11. The canonical Quantum Field Theory is widely believed to J. Foukzon et al. DOI: 10.4236/***.2019.***** 12 Journal of Modern Physics break down at some fundamental high-energy cutoff ∗Λ and therefore the quantum fluctuations in the vacuum can be treated classically seriously only up to this high-energy cutoff, see for example [15]. In this paper we argue that Quantum Field Theory in fractal space-time with negative Hausdorff-Colombeau dimensions [12] gives high-energy cutoff on natural way. 2. Ghosts as Physical Dark Matter 2.1. Paulu-Villars Ghosts As Physical Dark Matter Before explaining the role of PV ghosts, etc. as physical dark matter remind the idea of PV regularization as a conventional UV regularization. We consider, as an example, the scalar field theory with the interaction 4λφ . Lagrangian density of this theory reads 2 2 401 . 2 2 mμ μφ φ φ λφ= ∂ ∂ − +L (62) This theory requires UV regularization (e.g. in (2+1) and (3+1) dimensions). Let us show that it is sufficient to introduce N extra fields with large mass playing the role of the regularization parameter. Lagrangian density can be rewritten as follows ( ) 2 2 4 0 0 0 0 11 : :, 2 2 , . iN i ii N N i i ii i m a μ μφ φ φ λ φ φ φ φ φ φ φ = = =   = − ∂ ∂ − +    = + = = ∑ ∑ ∑  L (63) Here the symbol "::" means that in perturbation theory we drop Feynman diagrams with loops containing only one vertex. The 0φ is usual field with mass 0m and the , 1, ,i i Nφ =  is the extra field with mass , 1, ,im i N=  . It can be shown that in (3+1)-dimensional theory the introduction of one PV field is sufficient for the ultraviolet regularization of perturbation theory in λ . One can show that momentum space Feynman diagrams in the original theory with Lagrangian density (62) diverge no more than quadratically [16] [17] [18] (beside of vacuum diagrams) shown in Figure 1. If we consider now Feynman diagrams in the theory with Lagrangian density (63) we see that propagators of fields 0φ and φ  sum up in corresponding diagrams so that we obtain the following expression which plays the role of regularized propagator ( )2 2 2 2 2 2 20 0 0 1 , 0 0 0 N Nj j j j j j a a k k m i k m i k m i= = ∆ = = − − + − + − +∑ ∑ (64) where 2 2 2 2 20 1 2 3k k k k k= − + + . Integral corresponding to vacuum diagram is ( ) ( ) ( ) 4 4 2 4 4 2 20 d d . 02π 2π N j j j ak kk k m i= I = ∆ = − + ∑∫ ∫ (65) To do this integral, since it is convergent, we can Wick rotate. J. Foukzon et al. DOI: 10.4236/***.2019.***** 13 Journal of Modern Physics Figure 1. One-loop massive vacuum diagram. Remark 2.1.1. All the integrals in quantum field theory are written in Minkowski space, however, the ultraviolet divergence appears for large values of modulus of momentum and it is useful to regularise it in Euclidean space [17]. Transition to Euclidean space can be achieved by replacing thr zeroth component of momentum 0 4k ik→ , where the integration over the fourth component of momenta goes along the imaginary axis. To go to the integration along the real axis, one has to perform the (Wick) rotation of the integration contour by 90  (see Figure 2). This is possible since the integral over the big circle vanishes and during the transformation of the contour it does not cross the poles. Then we get 3 2 2 200 d . 8π N j E E E j E j a ki k k m ∞ = I = +∑∫ (66) To do this integral, since it is convergent, we can deal with regularized integral ( ) 3 2 2 20, d ,8π N j E E j E j a ki k k mε ε Λ = I Λ = + ∑∫ (67) where 0,ε Λ ∞  , i.e. ( ), EεI Λ ≈ I . We assume now that Pauli-Villars conditions given by Equations (48) holds. Let us consider now the quantity ( ) 3 2 2 20, d ,8π N j E E j E j a ki k k mη η ε ε η Λ = I I Λ = + ∑∫ (68) where ( ]0,1η ∈ , and therefore from Equation (68) we obtain 2 20 00 d d 0, 8π 8π N N E j E j E Ej j i ik a k a k kη ε εη Λ Λ = == I = = ≡∑ ∑∫ ∫ (69) since Equations (48) holds. From Equation (68) by differentiation we obtain ( ) 2 3 2 20 2 2 d d , d 8π N j j E E j E j a m ki k k m η εη η Λ = I = + ∑∫ (70) and therefore from Equation (39) we obtain J. Foukzon et al. DOI: 10.4236/***.2019.***** 14 Journal of Modern Physics Figure 2. The Wick rotation of the integration contour. ( ) 2 3 2 20 2 2 0 0 2 1 2 0 d d d 8π d 0, 8π N j j E E j E j N j j E Ej a m ki k k m i a m k k η ε η η ε η η Λ = = = Λ − = I = + = ≡ ∑∫ ∑ ∫ (71) since Equations (48) holds. From Equation (70) by differentiation we obtain ( ) ( ) ( ) ( ) 4 32 2 2 30 0 2 2 4 3 2 32 2 d d , d 4π d . 4π N N j j E j Ej j E j j j E j E E j a m ki k k m ia m kk k m η ε ε η η η η η Λ = = Λ I = R = + R = + ∑ ∑∫ ∫ (72) Note that ( ) ( ) 4 4 23 2 3 2 2 20 2 2 d . 4π 4π 4 16π j j j j j jE j E jE j ia m ia m a mk ik mk m η η ηη ∞ − R = = + ∫ (73) Thus ( ) 2 1 20 00 d d ln d 16π N N j j jj j a m η η η ηη = = I = R =∑ ∑∫ (74) and ( ) 2 20 ln ,16π N j j j a m η η η η=I = −∑ (75) Therefore ( ) 2 201 , 0, 16π N j j j a m η η ε == I Λ = I = − ≡∑ (76) since Equations (48) holds. Thus integral (65) corresponding to vacuum diagram by using Pauli-Villars renormalization identically equal zero, i.e. ( ) ( ) ( ) ( ) 4 4 2 4 4 2 20 d dRen 0. 02π 2π N j PV j j ak kk k m i= I = ∆ = ≡ − +∑ (77) J. Foukzon et al. DOI: 10.4236/***.2019.***** 15 Journal of Modern Physics Let us consider now how this method works in the case of the simplest scalar diagram shown in Figure 3. The corresponding Feinman integral has the form ( ) ( ) ( ) ( ) 4 2 4 2 2 2 2 2 0 0 1 d . 2π 0 0 kp k m i p k m i I =  − + − − +  ∫ (78) Regularized Feinman integral (78) reads ( ) ( ) ( ) ( ) 4 2 4 0 2 2 2 2 2 d1 , 2π 0 0 N j reg j j j a k p k m i p k m i= I =  − + − − +  ∑∫ (79) where 1N = . To do this integral, since it is convergent, we can Wick rotate. Then we get ( ) ( ) ( ) ( ) 4 2 4 0 2 2 2 2 2 d . 2π N j reg j j j a kip k m p k m= I =  + − +  ∑∫ (80) The integral (80) can be written as ( ) ( ) ( ) ( ) 41 2 4 20 2 2 2 0 31 2 20 2 2 2 0 d d 2π 1 d d . 8π 1 N j reg j j N j E E j E j a kip x k p x x m a k ki x k p x x m = = I =  + − +  =  + − +  ∑∫ ∫ ∑∫ ∫ (81) To do this integral, since it is convergent, we can deal with regularized integral ( ) ( ) 31 2 2 20 2 2 2 0 d , , d . 8π 1 N j E E reg j E j a k kip x k p x x mε ε Λ = I Λ =  + − +  ∑∫ ∫ (82) Let us consider now the quantity ( ) ( ) 31 2 2 20 2 2 2 0 d , , d . 8π 1 N j E E j E j a k kip x k p x x m η ε ε η Λ = I Λ =  + − +  ∑∫ ∫ (83) where ( ]0,1η ∈ , and therefore from Equation (83) we obtain ( )20 , , 0p εI Λ ≡ , since Equations (48) holds. From Equation (83) by differentiation we obtain Figure 3. The simplest scalar diagram. J. Foukzon et al. DOI: 10.4236/***.2019.***** 16 Journal of Modern Physics ( ) ( ) ( ) ( ) ( ) ( ) 2 31 2 2 30 2 2 2 0 2 2 2 0 31 2 32 2 2 0 1 2 2 0 dd , , d d 4π 1 , , , , 4π d , , , d 1 1 d . 4 1 N j j E E j E j N j j jj E E j E j j a m k kip x k p x x m i a m p k kp x k p x x m x p x x m η ε ε η η η ε η ε η η Λ = = I Λ = −  + − +  − R Λ R Λ  + − +  = − + ∑∫ ∫ ∑ ∫ ∫ ∫   (84) From Equation (84) we obtain ( ) ( ) ( ) 2 2 2 2 0 1 2 2 20 0 d , , , , , d 4π d . 16π 1 N j j jj N jj j ip a m p i xa m p x x η ε η εη η = −= I Λ − R Λ = − − + ∑ ∑ ∫  (85) From Equation (85) we obtain ( ) ( ) 1 1 2 2 2 20 0 0 dd . 16π 1 N reg jj j ip a x m p x x η η−= I = − − + ∑ ∫ ∫ (86) Note that ( ) ( ) ( ) ( ) ( ) ( ) ( ) 1 12 2 2 2 2 2 0 0 2 2 2 2 2 2 2 2 d 1 ln 1 1 1 1 1 ln 1 1 1 ln 1 1. j j j j j j j m p x x m p x x m p x x m p x x m p x x m p x x m p x x η η η η − − − − − − −    = − + − + −   − +    = − + − +       − − − −    ∫ (87) Thus ( ) ( ) ( ) ( ){ ( ) ( ) } ( ) ( ){ 1 1 12 2 2 20 0 0 1 1 2 2 2 2 2 0 0 12 2 2 2 2 0 1 1 2 2 2 2 2 0 0 dd 16π 1 d 1 1 ln 1 1 16π 1 ln 1 16π d 1 1 ln 1 1 16π N reg jj j N j j jj N j j jj N j j jj ip a x m p x x i a x m p x x m p x x im p x x m p x x a i a x m p x x m p x x η η = −= = − − = =− − = = − − = I = − − +    = − − + − +       − − − +       = − − + − +    ∑ ∫ ∫ ∑ ∫ ∑ ∑ ∫ ( ) ( ) }2 2 2 21 ln 1j jm p x x m p x x− −   − − −    ( ) ( ){ ( ) ( ) } ( ) ( ){ ( ) ( ) } 1 2 2 2 2 0 02 0 2 2 2 2 0 0 1 2 2 2 2 1 12 0 2 2 2 2 1 1 d 1 1 ln 1 1 16π 1 ln 1 d 1 1 ln 1 1 16π 1 ln 1 . i x m p x x m p x x m p x x m p x x i x m p x x m p x x m p x x m p x x − − − − − − − −    = − − + − +       − − −       + − + − +       − − −    ∫ ∫ (88) From Equation (88) we obtain J. Foukzon et al. DOI: 10.4236/***.2019.***** 17 Journal of Modern Physics ( ) ( ) ( ){ ( ) ( ) } ( ) ( ){ ( ) ( ) } 1 2 2 2 2 2 0 02 0 2 2 2 2 0 0 1 2 2 2 2 1 12 0 2 2 2 2 1 1 d 1 1 ln 1 1 16π 1 ln 1 d 1 1 ln 1 1 16π 1 ln 1 . reg ip x m p x x m p x x m p x x m p x x i x m p x x m p x x m p x x m p x x − − − − − − − −    I = − − + − +       − − −       + − + − +       − − −    ∫ ∫ (89) We assume now that 2 21 1m p −  and from Equation (89) finally we obtain ( ) ( ) ( ){ ( ) ( ) } ( ) 1 2 2 2 2 2 0 02 0 2 2 2 2 2 2 0 0 1 d 1 1 ln 1 1 16π 1 ln 1 . reg ip x m p x x m p x x m p x x m p x x O m p − − − − −    I = − − + − +       − − − +    ∫ (90) Remark 2.1.2. The simple renormalizable models with finite masses , 1, ,im i N=  which we have considered in the section many years regarded only as constructs for a study of the ultraviolet problem of QFT. The difficulties with unitarity appear to preclude their direct acceptability as canonical physical theories in locally Minkowski space-time. However, for their unphysical behavior may be restricted to arbitrarily large energy scales ∗Λ mentioned above by an appropriate limitation on the finite masses im . 2.2. Renormalizability of Higher Derivative Quantum Gravity Gravitational actions which include terms quadratic in the curvature tensor are renormalizable. The necessary Slavnov identities are derived from Becchi-Rouet-Stora (BRS) transformations of the gravitational and Faddeev-Popov ghost fields. In general, non-gauge-invariant divergences do arise, but they may be absorbed by nonlinear renormalizations of the gravitational and ghost fields and of the BRS transformations [14]. The geneic expression of the action reads ( )4 2 2d 2 ,symI x g R R R Rμνμνα β κ −= − − − +∫ (91) where the curvature tensor and the Ricci is defined by Rλ λμαν ν μα= ∂ Γ and R Rλμν μλν= correspondingly, 2 32πGκ = . The convenient definition of the gravitational field variable in terms of the contravariant metric density reads .h g gμν μν μνκ η= − − (92) Analysis of the linearized radiation shows that there are eight dynamical degrees of freedom in the field. Two of these excitations correspond to the familiar massless spin-2 graviton. Five more correspond to a massive spin-2 particle with mass 2m . The eighth corresponds to a massive scalar particle with mass 0m . Although the linearized field energy of the massless spin-2 and massive scalar excitations is positive definite, the linearized energy of the massive spin-2 excitations is negative definite. This feature is characteristic of higher-derivative models, and poses the major obstacle to their physical interpretation. In the quantum theory, there is an alternative problem which may be substituted for the negative energy. It is possible to recast the theory so that the masJ. Foukzon et al. DOI: 10.4236/***.2019.***** 18 Journal of Modern Physics sive spin-2 eigenstates of the free-fieid Hamiltonian have positive-definite energy, but also negative norm in the state vector space. These negative-norm states cannot be excluded from the physical sector of the vector space without destroying the unitarity of the S matrix. The requirement that the graviton propagator behaves like 4p− for large momenta makes it necessary to choose the indefinite-metric vector space over the negative-energy states. The presence of massive quantum states of negative norm which cancel some of the divergences due to the massless states is analogous to the Pauli-Villars regularization of other field theories. For quantum gravity, however, the resulting improvement in the ultraviolet behavior of the theory is sufficient only to make it renormalizable, but not finite. The gauge choice which we adopt in order to define the quantum theory is the canonical harmonic gauge: 0hμνν∂ = . Corresponding Green's functions are then given by a generating functional ( ) ( ) ( ) 4 4 4 d d d exp d d .sym Z T N h C C F i I xC F D C xT h μν σ τ μν τμ ν τ μν α μν τ μν α μν δ κ ≤      =      × + +  ∏∫ ∫ ∫  (93) Here , rF F h Fτ τ μν τμν μν μ νδ= = ∂   and the arrow indicates the direction in which the derivative acts. N is a normalization constant. Cσ is the Faddeev-Popov ghost field, and Cτ is the antighost field. Notice that both C σ and Cτ are anticommuting quantities. Dμνα is the operator which generates gauge transformations in hμν , given an arbitrary spacetime-dependent vector ( )xαξ corresponding to x xμ μ μκξ′ = + and where ( ) ( ) D x h h h h μν α μ ν ν μ μν α α α μ αν ν αμ α μν α μν α α α α ξ ξ ξ η ξ κ ξ ξ ξ ξ = ∂ + ∂ − ∂ + ∂ + ∂ − ∂ − ∂ (94) In the functional integral (93), we have written the metric for the gravitational field as dhμν μ ν≤    ∏ without any local factors of ( )detg gμν= . Such factors do not contribute to the Feynman rules because their effect is to introduce terms proportional to ( ) ( )4 40 d lnx gδ −∫ into the effective action and ( )4 0δ is set equal to zero in dimensional regularization. In calculating the generating functional (93) by using the loop expansion, one may represent the δ function which fixes the gauge as the limit of a Gaussian, discarding an infinite normalization constant ( )4 1 4 0 1lim exp d . 2 F i xF Fτ ττδ − ∆→   ∆     ∫ (95) In this expression, the index τ has been lowered using the flat-space metric tensor μνη . For the remainder of this paper, we shall adopt the standard approach to the covariant quantization of gravity, in which only Lorentz tensors occur, and all raising and lowering of indices is done with respect to flat space. J. Foukzon et al. DOI: 10.4236/***.2019.***** 19 Journal of Modern Physics The graviton propagator may be calculated from 1 41 d 2sym I xF Fττ −+ ∆ ∫ in the usual fashion, letting 0∆ → after inverting. The expression 1 41 d 2 xF Fττ −∆ ∫ contains only two derivatives. Consequently, there are parts of the graviton propagator which behave like 2p− for large momenta. Specifically, the 2p− terms consist of everything but those parts of the propagator which are transverse in all indices. These terms give rise to unpleasant infinities already at the one-loop order. For example, the graviton self-energy diagram shown in Figure 4 has a divergent part with the general structure ( )24h∂ . Such divergences do cancel when they are connected to tree diagrams whose outermost lines are on the mass shell, as they must if the S matrix is to be made finite without introducing counterterms for them. However, they greatly complicate the renormalization of Green's functions. We may attempt to extricate ourselves from the situation described in the last paragraph by picking a different weighting functional. Keeping in mind that we want no part of the graviton propagator to fall off slower than 4p− for large momenta, we now choose the weighting functional [12] ( ) 1 4 24 1exp d ,2e i xe e τ τ τω −  = ∆     ∫  (96) where eτ is any four-vector function. The corresponding gauge-fixing term in the effective action is 2 1 4 21 d . 2 xF Fττκ −− ∆ ∫  (97) The graviton propagator resulting from the gauge-fixing term (97) is derived in [13]. For most values of the parameters α and β in symI it satisfies the requirement that all its leading parts fall off like 4p− for large momenta. There are, however, specific choices of these parameters which must be avoided. If 0α = , the massive spin-2 excitations disappear, and inspection of the graviton propagator shows that some terms then behave like 2k − . Likewise, if 3 0β α− = , the massive scalar excitation disappears, and there are again terms in the propagator which behave like 2p− . However, even if we avoid the special cases 0α = and 3 0β α− = , and if we use the propagator derived from (97), we still do not obtain a clean renormalization of the Green's functions. We now turn to the implications of gauge invariance. Before we write down the BRS transformations for gravity, let us first establish the commutation relation for gravitational gauge transformations, which reveals the group structure of the theory. Take the gauge transformation (94) of hμν , generated by μξ and Figure 4. The one-loop graviton self-energy diagram. J. Foukzon et al. DOI: 10.4236/***.2019.***** 20 Journal of Modern Physics perform a second gauge transformation, generated by μη , on the hμν fields appearing there. Then antisymmetrize in μξ and μη . The result is ( ) ( ) ,D D D h μν ρσ α β α β μν λ α α λα β λ α αρσ δ ξ η η ξ κ ξ η ξ η δ − = ∂ − ∂ (98) where the repeated indices denote both summation over the discrete values of the indices and integration over the spacetime arguments of the functions or operators indexed. The BRS transformations for gravity appropriate for the gauge-fixing term (96) are [13] ( ) ( ) ( ) BRS 2 BRS 3 1 2 BRS a , b , c , h D C C C C C F μν μν α α α α β β τ τ δ κ δλ δ κ δλ δ κ δλ− = = − ∂ = − ∆  (99) where δλ is an infinitesimal anticommuting constant parameter. The importance of these transformations resides in the quantities which they leave invariant. Note that ( )BRS 0C Cσ ββδ ∂ = (100) and ( )BRS 0.D Cμν ααδ = (101) As a result of Equation (101), the only part of the ghost action which varies under the BRS transformations is the antighost Cτ . Accordingly, the transformation (99c) has been chosen to make the variation of the ghost action just cancel the variation of the gauge-fixing term. Therefore, the entire effective action is BRS invariant: 2 1 2 BRS 1 0. 2sim I F F C F D Cτ τ μν ατ τ μν αδ κ − − ∆ + =     (102) Equations (99), (100), and (102) now enable us to write the Slavnov identities in an economical way. In order to carry out the renormalization program, we will need to have Slavnov identities for the proper vertices. A) Slavnov identities for Green's functions First consider the Slavnov identities for Green's functions. ( ) ( ) , , , , d d d exp , , , , , . Z T K L N h C C i h C C K L C C C T h τ μν σ μν σ μν σ τμ ν μν σ σ σ τ μν τ μν σ σ σ τ μν β β β β β κ ≤      =      × + + +  ∏∫ Σ (103) Anticommuting sources have been included for the ghost and antighost fields, and the effective action Σ has been enlarged by the inclusion of BRS invariant couplings of the ghosts and gravitons to some external fields Kμν (anticommuting) and Lσ (commuting), 2 1 2 21 . 2sim I F F C F D C K D L C Cτ τ μν α μν σ βτ τ μν α μν α σ βκ κ κ −= − ∆ + + + ∂  Σ  (104) J. Foukzon et al. DOI: 10.4236/***.2019.***** 21 Journal of Modern Physics Σ is BRS invariant by virtue of Equation (99), Equation (100), and Equation (102). We may use the new couplings to write this invariance as 3 1 2 .F K L Ch C τμν σμν σ τ δ δ δ δ δκ δ δ δδ δ −+ + ∆     Σ Σ Σ Σ Σ  (105) In this equation, and throughout this subsection, we use left variational derivatives with respect to anticommuting quantities: ( )f C C f Cσ τ τδ δ δ δ= . Equation (105) may be simplified by rewriting it in terms of a reduced effective action, 2 1 21 . 2 F Fττκ −Σ = + ∆Σ  (106) Substitution of (106) into (105) gives 0, K Lh Cμν σμν σ δ δ δ δ δ δδ δ + = Σ Σ Σ Σ (107) where we have used the relation 1 0.F K C τ μν μν τ δ δκ δ δ − − =  Σ Σ (108) Note that a measure d d dh C Cμν σ τμ ν≤        ∏ (109) is BRS invariant since for infinitesimal transformations, the Jacobian is 1, because of the trace relations ( ) ( ) ( ) ( ) 2 2 a 0, b 0, K h C L μν μν σ σ δ δ δ δ δ δ = =   Σ Σ (110) both of which follow from 4d 0x Cαα∂ =∫ . The parentheses surrounding the indices in (110a) indicate that the summation is to be carried out only for μ ν≤ . Remark 2.2.1. Note that the Slavnov identity for the generating functional of Green's functions is obtained by performing the BRS transformations (99) on the integration variables in the generating functional (103). This transformation does not change the value of the generating functional and therefore we obtain ( ) ( ) 2 2 3 1 2 d d d exp 0. N h C C T D C C F h i T h C C μν σ μν σ β τ μν α σ βμ ν τ μν μν σ τ τμν μν σ τ κ κ β κ β κ β β ≤ −       − ∂     + ∆ + + + =  ∏∫  Σ (111) Another identity which we shall need is the ghost equation of motion. To derive this equation, we shift the antighost integration variable Cτ to C Cτ τδ+ , again with no resulting change in the value of the generating functional: J. Foukzon et al. DOI: 10.4236/***.2019.***** 22 Journal of Modern Physics ( ) d d d exp N h C C C i T h C C μν σ τ τ σμ ν μν σ τ μν σ τ δ β δ κ β β ≤        +        × + + +  ∏∫   Σ Σ (112) We define now the generating functional of connected Green's functions as the logarithm of the functional (103), , , , , ln , , , , .W T K L i Z T K Lτ τμν σ μν σ μν σ μν σβ β β β   = −    (113) and make use of the couplings to the external fields Kμν and Lσ to rewrite (112) in terms of W 2 1 2 0.W W WT F K L T τ μν σ τμν μν σ μν δ δ δκ β κ β δ δ δ −− + ∆ =   (114) Similarly, we get the ghost equation of motion: 1 0.WF K τ τ μν μν δκ β δ − + =  (115) B) Proper vertices A Legendre transformation takes us from the generating functional of connected Green's functions (113) to the generating functional of proper vertices. First, we define the expectation values of the gravitational, ghost, and antighost fields in the presence of the sources ,Tμν σβ , and τβ and the external fields Kμν and Lσ ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) a , b , c . Wh x T x WC x x WC x x μν μν σ σ τ τ δ κδ δ δβ δ δβ = = = (116) We have chosen to denote the expectation values of the fields by the same symbols which were used for the fields in the effective action (104). The Legendre transformation can now be performed, giving us the generating functional of proper vertices as a functional of the new variables (116) and the external fields Kμν and Lσ  , , , , , , , , . h C C K L W T K L T h C C μν σ τ μν σ τ μν σ τ μν σ μν σ μν σ τβ β κ β β  Γ    = − − −  (117) In this equation, the quantities Tμν , σβ , and τβ are given implicitly in terms of , , ,h C C Kμν σ τ μν , and Lσ by Equation (116). The relations dual to (116) are ( ) ( )  ( ) ( ) ( )  ( ) ( ) ( )  ( ) a , b , c . T x h x x C x x C x μν μν σ σ τ τ δκ δ δβ δ δβ δ Γ = − Γ = Γ = − (118) J. Foukzon et al. DOI: 10.4236/***.2019.***** 23 Journal of Modern Physics Since the external fields Kμν and Lσ do not participate in the Legendre transformation (116), for them we have the relations ( )  ( ) ( ) ( )  ( ) ( ) a , b . W K x K x W L x L x μν μν σ σ δ δ δ δ δ δ δ δ Γ = Γ = (119) Finally, the Slavnov identity for the generating functional of proper vertices is obtained by transcribing (114) using the relations (116), (118), and (119)      3 1 2 0.F h K Lh C C μν τμνμν σ σ μν σ δ δ δ δ δκ δ δδ δ δ −Γ Γ Γ Γ Γ+ + ∆ =   (120) We also have the ghost equation of motion,   1 0.F K C τ μν σ μν δ δκ δ δ − Γ Γ− =  (121) Since Equation (120) has exactly the same form as (105), we follow the example set by (106) and define a reduced generating functional of the proper vertices,  ( ) ( )2 1 21 .2 F h F h μν τ ρσ τμν ρσκ −Γ = Γ + ∆    (122) Substituting this into (120) and (121), the Slavnov identity becomes 0. K Lh Cμν σμν σ δ δ δ δ δ δδ δ Γ Γ Γ Γ + = (123) and the ghost equation of motion becomes 1 0.F K C τ μν μν τ δ δκ δ δ − Γ Γ− =  (124) Equations (123) and (124) are of exactly the same form as (107) and (108). This is as it should be, since at the zero-loop order ( )0 .Γ = Σ (125) C) Structure of the divergences and renormalization equation The Slavnov identity (123) is quadratic in the functional Γ . This nonlinearity is reflected in the fact that the renormalization of the effective action generally also involves the renormalization of the BRS transformations which must leave the effective action invariant. The canonical approach uses the Slavnov identity for the generating functional of proper vertices to derive a linear equation for the divergent parts of the proper vertices. This equation is then solved to display the structure of the divergences. From this structure, it can be seen how to renormalize the effective action so that it remains invariant under a renormalized set of BRS transformations [14]. Suppose that we have successfully renormalized the reduced effective action J. Foukzon et al. DOI: 10.4236/***.2019.***** 24 Journal of Modern Physics up to 1n − loop order; that is, suppose we have constructed a quantum extension of Σ which satisfies Equations (107) and (108) exactly, and which leads to finite proper vertices when calculated up to order 1n − . We will denote this renormalized quantity by ( )1n−Σ . In general, it contains terms of many different orders in the loop expansion, including orders greater than 1n − . The 1n − loop part of the reduced generating functional of proper vertices will be denoted by ( )1n−Γ . When we proceed to calculate ( )nΓ , we find that it contains divergences. Some of these come from n-loop Feynman integrals. Since all the subintegrals of an n-loop Feynman integral contain less than w loops, they are finite by assumption. Therefore, the divergences which arise from w-Ioop Feynman integrals come only from the overall divergences of the integrals, so the corresponding parts of ( )nΓ are local in structure. In the dimensional regularization procedure, these divergences are of order ( ) 11 4d −− = − , where d is the dimensionality of spacetime in the Feynman integrals. There may also be divergent parts of ( )nΓ which do not arise from loop integrals, and which contain higher-order poles in the regulating parameter  . Such divergences come from n-loop order parts of ( )1n−Σ which are necessary to ensure that (107) is satisfied. Consequently, they too have a local structure. We may separate the divergent and finite parts of ( )nΓ : ( ) ( ) ( ) div finite . n n nΓ = Γ + Γ (126) If we insert this breakup into Equation (123), and keep only the terms of the equation which are of n-loop order, we get ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 0 0 0 0 div div div div finite finite finite finite =0 . n n n n n i i n i i n i K K L Lh h C C K Lh C μν μν σ σ μν μν σ σ μν σ μν σ δ δ δ δδ δ δ δ δ δ δ δδ δ δ δ δ δ δ δ δ δδ δ − − Γ Γ Γ ΓΓ Γ Γ Γ + + +  Γ Γ Γ Γ = − +     ∑ (127) Since each term on the right-hand side of (127) remains finite as 0→ , while each term on the left-hand side contains a factor with at least a simple pole in e, each side of the equation must vanish separately. Remembering the Equation (125), we can write the following equation, called the renormalization equation: ( ) div 0, nRΓ = (128) where . K L K Lh C h Cμν σ μν σμν σ μν σ δ δ δ δ δ δ δ δ δ δ δ δδ δ δ δ Σ Σ Σ Σ R = + + + (129) Similarly by collecting the n-loop order divergences in the ghost equation of motion (124) we get ( ) ( ) 1 div div 0. n n F K C τ μν μν τ δ δ κ δ δ − Γ Γ− =  (130) J. Foukzon et al. DOI: 10.4236/***.2019.***** 25 Journal of Modern Physics In order to construct local solutions to Equations. (128) and (130) remind that the operator R defined in (129) is nilpotent [14]: 2 0.R = (131) Equation (131) gives us the local solutions to Equation (128) of the form ( ) ( ) ( )div , , , , ,n h X h C C K Lμν μν σ τ μν σ Γ = I +R  (132) where I is an arbitrary gauge-invariant local functional of hμν and its derivatives, and X is an arbitrary local functional of , , ,h C C Kμν σ τ μν and Lσ and their derivatives. In order to satisfy the ghost equation of motion (130) we require that ( ) ( ) ( )1div div , , , .n n h C K C F Lμν σ τμν τ μν σκ −Γ = Γ −  (133) D) Ghost number and power counting Structure of the effective action (104) shows that we may define the following conserved quantity, called ghost number [14]: [ ] [ ] ghost ghost ghost ghost ghost 0, 1, 1, 1, 2. N h N C N C N K N L μν σ τ μν σ    = = + = −     = − = −  (134) From Equations (134) follows that [ ] [ ]ghost ghost 0.N NΣ = Γ = (135) Since [ ]ghost 1,N R = + (136) we require of the functional ( )X ⋅ that [ ]ghost 1.N X = − (137) In order to complete analysis of the structure of ( )div nΓ , we must supplement the symmetry Equations (132), (133), and (137) with the constraints on the divergences which arise from power counting. Accordingly, we introduce the following notations: En = number of graviton vertices with two derivatives, Gn = number of antighost-graviton-ghost vertices, Kn = number of K-graviton-ghost vertices, Ln = number of L-ghost-ghost vertices, GI = number of internal-ghost propagators, CE = number of external ghosts, CE = number of external antighosts. Since graviton propagators behave like 4p− , and ghost propagators like 2p− , we are led by standard power counting to the degree of divergence of an arbitrary diagram, 4 2 2 2 3 3 .E G G L K CD n I n n n E= − + − − − − (138) The last term in (2.2.48) arises because each external antighost line carries with it a factor of external momentum. We can make use of the topological relation J. Foukzon et al. DOI: 10.4236/***.2019.***** 26 Journal of Modern Physics 2 2 2G G L K C CI n n n E E− = + − − (139) to write the degree of divergence as 4 2 2 2 .E L K C CD n n n E E= − − − − − (140) Together with conservation of ghost number, Equation (140) enables us to catalog three different types of divergent structures involving ghosts. These are illustrated in Figure 5. Each of the three types has degree of divergence 1 2 ED n= − . Consequently, all the divergences which involve ghosts have 0En = . Since the degree of divergence is then 1, the associated divergent structures in ( )div nΓ have an extra derivative appearing on one of the fields. Diagrams whose external lines are all gravitons have degree of divergence 4 2 ED n= − . Combining (140) with (137), (133), and (132), we can finally write the most general expression for ( )div nΓ which satisfies all the constraints of symmetries and power counting: ( ) ( ) ( ) ( ) ( )1div ,n h K C F P h L Q h Cμν τ μν αβ σ αβ τμν τ μν σ τκ − Γ = I +R − +   (141) where ( )P hμν αβ and ( )Q hσ αβτ are arbitrary Lorentz-covariant functions of the gravitational field hμν , but not of its derivatives, at a single spacetime point. ( )hμνI is a local gauge-invariant functional of hμν containing terms with four, two, and zero derivatives. Expanding (141), we obtain an array of possible divergent structures: ( ) ( ) ( ) ( ) ( ) ( ) ( ) div 2 2 2 . symn I Dh P K C F C P h h PK C F D C K C F D Q C h L Q C C L C Q C Q L C D C L Q C C h ρσ μν μν τ α μνα ρσ τ ρσμν μν ρσ τ μν α τ μν σ ε ρσ τ ρσ α μν τ μν σ εμν σ τ β σ β τ σ β τ σ β τ σ τ μν α σ τ βτ σ α σ τ βμν δ δ κ δ δ δκ κ δ κ κ δ κ κ δ   Γ = I + + −     − − − − − ∂ − ∂ − + ∂    (142) is determined only up to a term of the form [13] [14] Figure 5. The three types of divergent diagram which involve external ghost lines. Arbitrarily many gravitons may emerge from each of the central regions,(a) Ghost action type,(b) K type, (c) L type. J. Foukzon et al. DOI: 10.4236/***.2019.***** 27 Journal of Modern Physics The breakup between the gauge-invariant divergences S and the rest of (142) ( )4d ,symIx h h μν μν μν δ η κ κδ +∫ (143) which can be generated by adding to Pμν a term proportional to h g gμν μν μνη κ+ = . The profusion of divergences allowed by (142) appears to make the task of renormalizing the effective action rather complicated. Although the many divergent structures do pose a considerable nuisance for practical calculations, the situation is still reminiscent in principle of the renormalization of Yang-Mills theories. There, the non-gauge-invariant divergences may be eliminated by a number of field renormalizations. We shall find the same to be true here, but because the gravitational field hμν carries no weight in the power counting, there is nothing to prevent the field renormalizations from being nonlinear, or from mixing the gravitational and ghost fields. The corresponding renormalizations procedure considered in [13] [14]. Remark 2.2.2. We assume now that: 1) The local Poincaré group of momentum space is deformed at some fundamental high-energy cutoff ∗Λ [9] [10]. 2) The canonical quadratic invariant 2 ab a bp p pη= collapses at high-energy cutoff ∗Λ and being replaced by the non-quadratic invariant: ( ) 2 0 . 1 ab a bp pp l p η ∗Λ = + (144) 3) The canonical concept of Minkowski space-time collapses at a small distance 1l ∗ − Λ ∗= Λ to fractal space-time with Hausdorff-Colombeau negative dimension and therefore the canonical Lebesgue measure 4d x being replaced by the Colombeau-Stieltjes measure ( )( ) ( )( )( )4d , d ,x v s x xεε εη ε = (145) where ( )( )( ) ( ) ( ) 1 , , D v s x s x s x x x ε ε ε μ μ ε − −  = +      = (146) see subsection IV.2. 4) The canonical concept of local momentum space collapses at fundamental high-energy cutoff ∗Λ to fractal momentum space with Hausdorff-Colombeau negative dimension and therefore the canonical Lebesgue measure 3d k , where ( ), ,x y zk k k=k being replaced by the Hausdorff-Colombeau measure ( ) ( ) ( ) 1, d dd , D D D D D D D D D p p p ε ε ε ε + + + − − − − + − −∆ ∆ ∆ =    + +        k k k (147) see Subsection 3.4. Note that the integral over measure ,dD D + − k is given by J. Foukzon et al. DOI: 10.4236/***.2019.***** 28 Journal of Modern Physics formula (185). Remark 2.2.3. (I) The renormalizable models which we have considered in this section many years regarded only as constructs for a study of the ultraviolet problem of quantum gravity. The difficulties with unitarity appear to preclude their direct acceptability as canonical physical theories in locally Minkowski space-time. In canonical case they do have only some promise as phenomenological models. (II) However, for their unphysical behavior may be restricted to arbitrarily large energy scales ∗Λ mentioned above by an appropriate limitation on the renormalized masses 2m and 0m . Actually, it is only the massive spin-two excitations of the field which give the trouble with unitarity and thus require a very large mass. The limit on the mass 0m is determined only by the observational constraints on the static field. 3. Hausdorff-Colombeau Measure and Associated Negative Hausdorff-Colombeau Dimension 3.1. Fractional Integration in Negative Dimensions Let DHμ + be a Hausdorff measure [19] and nX ⊂  is measurable set. Let ( )s x be a function :s X →  such that is symmetric with respect to some centre 0x X∈ , i.e. ( )s x = constant for all x satisfying ( )0,d x x r= for arbitrary values of r. Then the integral in respect to Hausdorff measure over n-dimensional metric space X is then given by [19]: ( ) ( ) ( ) 2 1 0 2πd d . 2 D D D HX s x s r r r D μ + + + ∞ − + = Γ∫ ∫ (148) The integral in RHS of the Equation (148) is known in the theory of the Weyl fractional calculus where, the Weyl fractional integral ( )DW f x , is given by ( ) ( ) ( ) ( ) 1 0 1 d .DDW f x t x f t t D ∞ −= − Γ ∫ (149) Remark 3.1.1. In order to extend the Weyl fractional integral (148) in negative dimensions we apply the Colombeau generalized functions [20] and Colombeau generalized numbers [21]. Recall that Colombeau algebras ( )Ω of the Colombeau generalized functions defined as follows. Let Ω be an open subset of n . Throughout this paper, for elements of the space ( )( ]0,1C∞ Ω of sequences of smooth functions indexed by ( ]0,1ε ∈ we shall use the canonical notations ( )( )xε εζ and ( )uε ε so ( )u Cε ∞∈ Ω , ( ]0,1ε ∈ . Definition 3.1.1. We set ( ) ( ) ( )/MΩ = Ω Ω   , where ( ) ( ) ( )( ]{ ( ) ( ) } ( ) ( ) ( )( ]{ ( ) ( ) } 0,1 0,1 , with sup as 0 , , sup as 0 . n M p x K n q x K u C K p u x O u C K q u x O ε ε ε ε ε ε α ε ε α ε ε ∞ − ∈ ∞ ∈ Ω = ∈ Ω ∀ ⊂⊂ Ω ∀ ∈ ∃ ∈ = → Ω = ∈ Ω ∀ ⊂⊂ Ω ∀ ∈ ∀ ∈ = →       (150) J. Foukzon et al. DOI: 10.4236/***.2019.***** 29 Journal of Modern Physics Notice that ( )Ω is a differential algebra. Equivalence classes of sequences ( )uε ε will be denoted by ( )cl uε ε   is a differential algebra containing ( )D′ Ω as a linear subspace and ( )C∞ Ω as subalgebra. Definition 3.1.2. Weyl fractional integral ( )DW f xε ε − −      in negative dimensions 0D− < , 0, 1, , , ,D n n− ≠ − − ∈   is given by ( ) ( ) ( ) ( )( ) ( )( ) ( ) ( ) ( ) 1 0 1 1 d or 1 1 d , DD D D W f x t x f t t D W f x f t t D t x ε ε ε ε ε ε −− − − − ∞ − − ∞ − + = − Γ    =  Γ + −  ∫ ∫ (151) where ( ]0,1ε ∈ and ( ) 0 df t t ∞ < ∞∫ . Note that ( )( ) ( )DW f xε ε −− ∈  . Thus in order to obtain appropriate extension of the Weyl fractional integral ( )DW f x+ on the negative dimensions 0D− < the notion of the Colombeau generalized functions is essentially important. Remark 3.1.2. Thus in negative dimensions from Equation (148) we formally obtain ( )( ) ( ) ( ) ( ) 2 , 0 1 2πd d , 2 D DD HCX D s r s x r I D r ε ε εε ε μ ε − −− − − ∞ − +    = =  Γ +  ∫ ∫ (152) where ( ]0,1ε ∈ and 0, 2, , 2 , ,D n n− ≠ − − ∈   and where ( ),DHC ε ε μ − is appropriate generalized Colombeau outer measure. Namely Hausdorff-Colombeau outer measure. Remark 3.1.3. Note that: if ( )0 0s ≠ the quantity ( ),D DIε ε + − takes infinite large value in sense of Colombeau generalized numbers, i.e., ( ),D DIε ε + − = ∞    , see Definition 3.3.2 and Definition 3.3.3. Remark 3.1.4. We apply through this paper more general definition then (3.1.4): ( )( ) ( ) ( ) ( ) ( ) 12 2 ,, , 0 4π πd d , 2 2 DD D D DD D HCX D r s r s x r I D D r ε ε ε ε ε μ ε ++ − ++ − − − − ∞ + −    = =  Γ Γ +  ∫ ∫ (153) where ( ]0,1ε ∈ and 1D+ ≥ , 0, 2, , 2 , ,D n n− ≠ − − ∈   and where ( ),,D DHC ε ε μ + − is appropriate generalized Colombeau outer measure. Namely Hausdorff-Colombeau outer measure. In Subsection 3.3 we pointed out that there exists Colombeau generalized measure ( ),,d D DHC ε ε μ + − and therefore Equation (151) gives appropriate extension of the Equation (148) on the negative Hausdorff-Colombeau dimensions. 3.2. Hausdorff Measure and Associated Positive Hausdorff Dimension Recall that the classical Hausdorff measure [19] [22] originate in Caratheodory's construction, which is defined as follows: for each metric space X, each set { }i iF E ∈=  of subsets iE of X, and each positive function ( )Eζ + , such that J. Foukzon et al. DOI: 10.4236/***.2019.***** 30 Journal of Modern Physics ( )0 iEζ +≤ ≤ ∞ whenever iE F∈ , a preliminary measure ( )Eδφ+ can be constructed corresponding to 0 δ< ≤ +∞ , and then a final measure ( )Eμ+ , as follows: for every subset E X⊂ , the preliminary measure ( )Eδφ+ is defined by ( ) { } ( ) ( ){ }inf | , . i i i i ii iE E E E E diam Eδφ ζ δ ∈ + + ∈ ∈ = ⊂ ≤∑     (154) Since ( ) ( ) 1 2 E Eδ δφ φ + +≥ for 1 20 δ δ< < ≤ +∞ , the limit ( ) ( ) ( ) 0 0 lim supE E Eδ δδ δ μ φ φ + + + + → > = = (155) exists for all E X⊂ . In this context, ( )Eμ+ can be called the result of Caratheodory's construction from ( )Eζ + on F. ( )Eδφ+ can be referred to as the size δ approximating positive measure. Let ( ),iE dζ + + be for example ( ) ( ) ( ) ( ), ,0 ,di iE d d diam E dζ + + + + + = Θ < Θ  (156) for non-empty subsets ,iE i∈ of X. Where ( )d +Θ is some geometrical factor, depends on the geometry of the sets iE , used for covering. When F is the set of all non-empty subsets of X, the resulting measure ( ),H E dμ+ + is called the d+-dimensional Hausdorff measure over X; in particular, when F is the set of all (closed or open) balls in X, ( ) ( ) ( )2π 2 1 .2 d d dd d + + + + + −  Θ Ω = Γ +     (157) Consider a measurable metric space ( )( ) ( ), , , ,nHX d X dμ ∈ −∞ +∞ . The elements of X are denoted by , , ,x y z  , and represented by n-tuples of real numbers ( )1 2, , , nx x x x=  The metric ( ),d x y is a function :d X X R+× → is defined in n dimensions by ( ) ( )( ) 1 2, ij i i j j ij d x y y x y xδ = − − ∑ (158) and the diameter of a subset E X⊂ is defined by ( ) ( ){ }sup , | , .diam E d x y x y E= ∈ (159) Definition 3.2.1. The Hausdorff measure ( ),H E Dμ+ + of a subset E X⊂ with the associated Hausdorff positive dimension D+ +∈ is defined by canonical way ( ) { } ( ) ( )( ){ } ( ) ( ){ } 0 , lim inf , | , , sup | 0, , . i i H i i ii iE H E D E D E E i diam E D E d d E d δ μ ζ δ μ ∈ + + + + ∈→ + + + + + +  = ⊂ ∀ <   = ∈ > = +∞ ∑     (160) Definition 3.2.2. Remind that a function :f X →  defined in a measurable space ( ), ,X μΣ , is called a simple function if there is a finite disjoint set of sets { }1, , , nE E of measurable sets and a finite set { }1, , nα α of real numbers such that ( ) if x α= if ix E∈ and ( ) 0f x = if ix E∉ . Thus J. Foukzon et al. DOI: 10.4236/***.2019.***** 31 Journal of Modern Physics ( ) ( )1 i n i Eif x xα χ== ∑ , where ( )iE xχ is the characteristic function of iE . A simple function f on a measurable space ( ), ,X μΣ is integrable if ( )iEμ < ∞ for every index i for which 0iα ≠ . The Lebesgue-Stieltjes integral of f is defined by ( )=1d . n i iif Eμ α μ= ∑∫ (161) A continuous function is a function for which ( ) ( )limx y f x f y→ = whenever ( )lim , 0x y d x y→ = . The Lebesgue-Stieltjes integral over continuous functions can be defined as the limit of infinitesimal covering diameter: when { }i iE ∈ is a disjoined covering and i ix E∈ by definition (3.2.12) one obtains ( ) ( ) ( ) ( ) { } ( )( )=0 with d , lim inf , . ii ij ij ij HX i ij ijE X jdiam E E E E f x x D f x E D E μ ζ + + + + → ⊃   =     ∫ ∑ ∑   (162) From now on, X is assumed metrically unbounded, i.e. for every x X∈ and 0r > there exists a point y such that ( ),d x y r> . The standard assumption that D+ is uniquely defined in all subsets E of X requires X to be regular (homogeneous, uniform) with respect to the measure, i.e. ( )( ) ( )( ), ,H r H rB x D B y Dμ μ+ + + += for all elements ,x y X∈ and (convex) balls ( )rB x and ( )rB y of the form ( ) ( ){ }0 | , ; ,rB x z d x z r x z X> = ≤ ∈ . In the limit ( ) 0idiam E → , the infimum is satisfied by the requirement that the variation overall coverings { }ij ijE ∈ is replaced by one single covering iE , such that ij i ij E E x→  . Hence ( ) ( ) ( ) ( ) ( ) 0 d , lim , . i i H i iX diam E E X f x x D f x E Dμ ζ+ + + + → = = ∑∫  (163) The range of integration X may be parametrized by polar coordinates with ( ),0r d x= and angle Ω . { },i ir iE Ω ∈ can be thought of as spherically symmetric covering around a centre at the origin. In the limit, the function ( ), ,rE Dζ + +Ω defined by Equation (156) is given by ( ) ( ) ( ) , 1 1 , 0 d , lim , d d . r D D H r diam E x D E D r r ω μ ζ + ++ + + + − − Ω → = = Ω (164) Let us assume now for simplification that ( ) ( ) ( )f x f x f r= = and ( )lim 0 r f r →∞ = . The integral over a D+ -dimensional metric space X is then given by ( ) ( ) ( ) ( ) 2 1 0 2πd , d d . 1 2 D D D HX X f x x D f x x f r r r D μ + + +∞+ + − + = =   Γ +    ∫ ∫ ∫ (165) The integral defined in (163) satisfies the following conditions. 1) Linearity: ( ) ( ) ( ) ( ) ( ) ( ) ( ) 1 1 2 2 1 1 2 2 d , d , d , . HX H HX X a f x a f x x D a f x x D a f x x D μ μ μ + + + + + + +   = + ∫ ∫ ∫ (166) J. Foukzon et al. DOI: 10.4236/***.2019.***** 32 Journal of Modern Physics 2) Translational invariance: ( ) ( ) ( ) ( )0 d , d ,H HX Xf x x x D f x x Dμ μ+ + + ++ =∫ ∫ (167) since ( ) ( )0d , d ,H Hx x D x Dμ μ+ + + +− = . 3) Scaling property: ( ) ( ) ( ) ( )d , d ,DH HX Xf ax x D a f x x Dμ μ ++ + − + +=∫ ∫ (168) since ( ) ( )d , d ,DH Hx a D a x Dμ μ++ + − + += . 4) The generalized ( )D xδ + function: The generalized ( )D xδ + function for sets with non-integer Hausdorff dimension exists and can be defined by formula ( ) ( ) ( ) ( )0 0d , .D HX f x x x x D f xδ μ + + +− =∫ (169) 3.3. Hausdorff-Colombeau Measure and Associated Negative Hausdorff-Colombeau Dimensions During last 20 years the notion of negative dimension in geometry was many developed, see [12] [23] [24] [25] [26] [27]. Remind that canonical definitions of noninteger positive dimension always equipped with a measure. Hausdorff-Besicovich dimension equipped with Hausdorff measure ( )d ,H x Dμ+ + . Let us consider example of a space of noninteger positive dimension equipped with the Haar measure. On the closed interval 0 1x≤ ≤ there is a scale 0 1σ≤ ≤ of Cantor dust with the Haar measure equal to xσ for any interval ( )0, x similar to the entire given set of the Cantor dust. The direct product of this scale by the Euclidean cube of dimension 1k − gives the entire scale k σ+ , where k ∈ and ( )0,1σ ∈ [24]. In this subsection we define generalized Hausdorff-Colombeau measure. In subsection III.4 we will prove that negative dimensions of fractal equipped with the Hausdorff-Colombeau measure in natural way. Let Ω be an open subset of n , let X be metric space nX  and let F be a set { }i iF E ∈=  of subsets iE of X. Let ( ), ,E x xζ  be a function : Fζ ×Ω×Ω→  . Let ( )FC∞ Ω be a set of the all functions ( ),E xζ such that ( ) ( ),E x Cζ ∞∈ Ω whenever E F∈ . Throughout this paper, for elements of the space ( )( ]0,1FC∞ Ω of sequences of smooth functions indexed by ( ]0,1ε ∈ we shall use the canonical notations ( )( ),E xε εζ and ( )ε εζ so ( )FCεζ ∞∈ Ω , ( ]0,1ε ∈ . Definition 3.3.1. We set ( ) ( ) ( ), / ,F M F FΩ = Ω Ω   , where ( ) ( ) ( )( ]{ ( ) ( ) } ( ) ( ) ( )( ]{ ( ) ( ) } 0,1 ; 0,1 ; , , with sup , as 0 , , , sup , as 0 . n M F p E F x K n F q E F x K F C K p E x O F C K q E x O ε ε ε ε ε ε ζ α ζ ε ε ζ α ζ ε ε ∞ − ∈ ∈ ∞ ∈ ∈ Ω = ∈ Ω ∀ ⊂⊂ Ω ∀ ∈ ∃ ∈ = → Ω − ∈ Ω ∀ ⊂⊂ Ω ∀ ∈ ∀ ∈ = →       (170) J. Foukzon et al. DOI: 10.4236/***.2019.***** 33 Journal of Modern Physics Notice that ( )F Ω is a differential algebra. Equivalence classes of sequences ( )ε εζ will be denoted by ( )cl ε εζ   or simply ( )ε εζ   . Definition 3.3.2. We denote by  the ring of real, Colombeau generalized numbers. Recall that by definition ( ) ( )/M=    [21], where ( ) ( ) ( ] ( ) ( ]( ){ } ( ) ( ) ( ] ( ) ( ]( ){ } 0,1 0 0 0,1 0 0 0,1 , 0,1 . M x x x x α ε εε α ε εε α ε ε ε ε α ε ε ε ε  = ∈ ∃ ∈ ∃ ∈ ∀ ≤ ≤   = ∈ ∀ ∈ ∃ ∈ ∀ ≤ ≤          (171) Notice that the ring  arises naturally as the ring of constants of the Colombeau algebras ( )Ω . Recall that there exists natural embedding :r    such that for all r∈ , ( )r rε ε= where r rε ≡ for all ( ]0,1ε ∈ . We say that r is standard number and abbreviate r∈ for short. The ring  can be endowed with the structure of a partially ordered ring: for ,r s∈  , 1η η+∈ ≤ we abbreviate ,r sη≤  or simply r s≤ if and only if there are representatives ( )rε ε and ( )sε ε with r sε ε≤ for all ( ]0,ε η∈ . Colombeau generalized number r∈  with representative ( )rε ε we abbreviate ( )cl rε ε   . Definition 3.3.3. 1) Let δ ∈  . We say that δ is infinite small Colombeau generalized number and abbreviate 0δ ≈    if there exists representative ( )ε εδ and some q∈ such that ( )qOεδ ε= as 0ε → . 2) Let ∆∈  . We say that ∆ is infinite large Colombeau generalized number and abbreviate ∆ = ∞    if 1 0−∆ ≈      . 3) Let ∞ be { }∞ We say that ∞Θ∈  is infinite Colombeau generalized number and abbreviate Θ = ∞     if there exists representative ( )ε εΘ where εΘ = ∞ for all ( ]0,1ε ∈ . Here we set ( ) ( ) ( ){ }M M ε ε∞ = Θ   , ( ) ( )∞ =   and ( ) ( )/M∞ ∞ ∞=    . Definition 3.3.4. The singular Hausdorff-Colombeau measure originate in Colombeau generalization of canonical Caratheodory's construction, which is defined as follows: for each metric space X, each set { }i iF E ∈=  of subsets iE of X, and each Colombeau generalized function ( )( ), ,E x xε εζ  , such that: 1) ( )( )0 , ,E x xε εζ≤  , 2) ( )( ), ,E x xε εζ = ∞    , whenever E F∈ , a preliminary Colombeau measure ( )( ), , ,E x xδ εφ ε  can be constructed corresponding to 0 δ< ≤ +∞ , and then a final Colombeau measure ( )( ), ,E x xε εμ  , as follows: for every subset E X⊂ , the preliminary Colombeau measure ( )( ), , ,E x xδ εφ ε  is defined by ( ) { } ( ) ( ){ }, , , sup , , | , . i i i i ii i E E x x E x x E E diam Eδ εφ ε ζ δ ∈ ∈ ∈ = ⊂ ≤∑      (172) Since for all ( ]0,1ε ∈ : ( ) ( )1 2, , , , , ,E x x E x xδ δφ ε φ ε − −≥  for 1 20 δ δ< < ≤ +∞ , the limit ( )( ) ( )( ) ( )( )0 0, , , lim , , , inf , , ,E x x E x x E x xδ δε δ δ εεμ ε φ ε φ ε+→ >= =   (173) exists for all E X⊂ . In this context, ( )( ), , ,E x x εμ ε  can be called the result of Caratheodory's construction from ( )( ), ,E x xε εζ  on F and ( )( ), , ,E x xδ εφ ε  can be referred to as the size δ approximating Colombeau measure. J. Foukzon et al. DOI: 10.4236/***.2019.***** 34 Journal of Modern Physics Definition 3.3.5. Let ( )( ), , , ,iE D D x xε εζ + −  be ( )( ) ( ) ( ) ( ) ( ) 1 2 if , , , , , 0 if D i iD i i D D diam E x E E D D x x d x x x E ε ε ε ζ ε + − + − + −   Θ Θ    ∈ =  +     ∉   (174) where ( ] ( ) ( )1 20,1 , , 0D Dε + −∈ Θ Θ > . In particular, when F is the set of all (closed or open) balls in X, ( ) 2 1 2 π 1 2 D D D D + +− + + Θ =   Γ +    (175) and ( ) ( ) 2 2 2 π , 1 2 2, 4, 6, , 2 1 , D D D D D n − −− − − − Θ =   Γ +    ≠ − − − − +  (176) Definition 3.3.6. The Hausdorff-Colombeau singular measure ( )( ), , , , ,H E D D x x εμ ε+ −  of a subset E X⊂ with the associated Hausdorff-Colombeau dimension ( ) ,D D D+ − −+ +∈ ∈    is defined by ( )( ) { } ( )( ) ( )( ){ } ( )( ) { } 0 , , , , , lim sup , , , , | , , sup 0 | , , , , , , i i HC i i ii i E HC E D D x x E D D x x E E i diam E D D E D D x x ε εδ ε ε ε μ ε ζ δ μ ε ∈ + − + − ∈→ + + + −    = ⊂ ∀ <      = > = ∞ ∑           (177) The Colombeau-Lebesgue-Stieltjes integral over continuous functions :f X →  can be evaluated similarly as in Subsection III.3, (but using the limit in sense of Colombeau generalized functions) of infinitesimal covering diameter when { }i iE ∈ is a disjoined covering and i ix E∈ : ( ) ( )( ) ( ) ( ) { } ( ) 0 with d , , , , , lim inf , , , , . ii ij ij ij HCX i i iE X jdiam E E E E f x E D D x x f x E D D x x ε ε ε μ ε ζ + − + − =→ ⊃    =       ∫ ∑ ∑     (178) We assume now that X is metrically unbounded, i.e. for every x X∈ and 0r > there exists a point y such that ( ),d x y r> . The standard assumption that D+  and D−  is uniquely defined in all subsets E of X requires X to be regular (homogeneous, uniform) with respect to the measure, i.e. ( )( )( ) ( )( )( ), , , , , , , , , ,HC r HC rB x D D x x B y D D x yε εμ ε μ ε− + − − + − ′=         , where ( ) ( ), ,d x x d x y′=  for all elements , , ,x y x x X′∈  and convex balls ( )rB x  and ( )rB y  of the form ( ) ( ){ }| , ; ,rB x z d x z r x z X= ≤ ∈   and J. Foukzon et al. DOI: 10.4236/***.2019.***** 35 Journal of Modern Physics ( ) ( ){ }| , ; ,rB y z d y z r y z X= ≤ ∈   . In the limit ( ) 0idiam E → , the infimum is satisfied by the requirement that the variation over all coverings { }ij ijE ∈ is replaced by one single covering iE , such that ij i ij E E x→  . Therefore ( ) ( )( ) ( ) ( ) ( ) 0 d , , , , , lim , , , , . ii HCX i i iE Xdiam E f x E D D x x f x E D D x x ε ε ε μ ε ζ + − + − =→  =     ∫ ∑        (179) Assume that ( ) ( ) ,f x f r r r= = . The range of integration X may be parametrized by polar coordinates with ( ),0r d x= and angle ω . { },i irE ω can be thought of as spherically symmetric covering around a centre at the origin. Thus ( ) ( )( ) ( ) ( ) ( ) 0 d , , , , , lim , , , , . ii HCX i i iE Xdiam E f r E D D x x f r E D D x x ε ε ε μ ε ζ + − + − =→  =     ∫ ∑        (180) Notice that the metric set nX ⊂  can be tesselated into regular polyhedra; in particular it is always possible to divide n into parallelepipeds of the form ( ) ( ){ }1 , , 1, , | 1 ,0 , 1, , .ni i n j j j j j jx x x X x i x x j nγ γΠ = = ∈ = − ∆ + ≤ ≤ ∆ =   (181) For 2n = the polyhedra 1 2,i i Π is shown in Figure 6. Since X is uniform ( )( ) ( ) ( ) ( ) 1 , ,1 , ,1 , , 1 1 d , , , , , lim , , , , lim d . n i in i in HC i i diam D n n j j Ddiam j j D n n j j D nD j j x D D x x D D x x x x x x x x εε ε ε ε μ ε ζ ε ε + − + + − + − + − Π =Π =   = Π        ∆  =     − +                  − +     ∏ ∏                  (182) Notice that the range of integration X may also be parametrized by polar coordinates with ( ),0r d x= and angle Ω . ,rE Ω can be thought of as spherically symmetric covering around a centre at the origin (see Figure 7 for the two-dimensional case). In the limit, the Colombeau generalized function ( )( ), , , , ,0rE D D rε εζ + −Ω   is given by ( )( ) ( ) { }( ) , ,1 1 1 , d , , , , d dlim , , , , ,0 i in HC D D r Ddiam r D D r rE D D r r ε ε ε ε μ ε ζ ε + + − + − − − + − Ω Π Ω   Ω = Ω     +             (183) J. Foukzon et al. DOI: 10.4236/***.2019.***** 36 Journal of Modern Physics Figure 6. The polyhedra covering for 2n = . Figure 7. The spherical covering ,rE Ω . When ( )f x is symmetric with respect to some centre x X∈ , i.e. ( )f x = constant for all x satisfying ( ),d x x r= for arbitrary values of r, then change of the variable x z x x→ = −  (184) can be performed to shift the centre of symmetry to the origin (since X is not a linear space, (184) need not be a map of X onto itself and (184) is measure presuming). The integral over metric space X is then given by formula ( ) ( )( ) ( ) ( ) ( )12 2 0 d , , , , , 4π π d . 2 2 HCX DD D D f x E D D x x r f r r D D r ε ε μ ε ε ++ − − + − − ∞ + −    =  Γ Γ +  ∫ ∫    (185) 3.4. Main Properties of the Hausdorff-Colombeau Metric Measures with Associated Negative Hausdorff-Colombeau Dimensions Definition 3.4.1. An outer Colombeau metric measure on a set nX  is a Colombeau generalized function ( )( ) ( )FEε εφ  ∈ Ω   (see Definition 3.3.1) defined on all subsets of X satisfies the following properties. 1) Null empty set: The empty set has zero Colombeau outer measure J. Foukzon et al. DOI: 10.4236/***.2019.***** 37 Journal of Modern Physics ( )( ) 0.ε εφ ∅ =  (186) 2) Monotonicity: For any two subsets A and B of X ( )( )  ( )( ) .A B A Bε εε εφ φ   ≤    (187) 3) Countable subadditivity: For any sequence { }jA of subsets of X pairwise disjoint or not ( )( )  ( )( )11 .j jjj A Aε ε εεφ φ ∞ ∞ ==    ≤       ∑  (188) 4) Whenever ( ) ( ){ }, inf , | , 0nd A B d x y x A y B= ∈ ∈ > ( )( ) ( )( ) ( )( ) ,A B A Bε ε εε ε εφ φ φ     = +      (189) where ( ),nd x y is the usual Euclidean metric: ( ) ( ) 2,n i id x y x y= −∑ . Definition 3.4.2. We say that outer Colombeau metric measure ( ) ( ], 0,1ε εμ ε ∈ is a Colombeau measure on σ-algebra of subests of nX  if ( )ε εμ satisfies the following property: ( )( ) ( )( )11 .j jjj A Aε ε εεφ φ ∞ ∞ ==    =       ∑ (190) Definition 3.4.3. If U is any non-empty subset of n-dimensional Euclidean space, n , the diamater U of U is defined as ( ){ }sup , | ,U d x y x y U= ∈ (191) If nF ⊆  , and a collection { }i iU ∈ satisfies the following conditions: 1) iU δ< for all i∈ , 2) iiF U∈⊆  , then we say the collection { }i iU ∈ is a δ-cover of F. Definition 3.4.4. If nF ⊆  and , , 0s q δ > , we define Hausdorff-Colombeau content: ( )( ), 1, inf s is q qi i U H F x δ ε ε ε ε ∞ =     =   +    ∑ (192) where the infimum is taken over all δ-covers of F and where ( ),1 ,, ,i i i n ix x x U= ∈ for all i∈ and x is the usual Euclidean norm: 2 1 n jjx x== ∑ . Note that for ( ]1 20 1, 0,1δ δ ε< < < ∈ we have ( ) ( ) 1 2 , ,, ,s q s qH F H Fδ δε ε≥ (193) since any 1δ cover of F is also a 2δ cover of F, i.e. ( )1 , ,s qH Fδ ε is increasing as δ decreases. Definition 3.4.5. We define the ( ),s q -dimensional Hausdorff-Colombeau (outer) measure as: ( )( ) ( )( ), ,0, lim , .s q s qH F H Fδε δ εε ε→= (194) Theorem 3.4.1. For any δ-cover, { }i iU ∈ of F, and for any ( ]0,1ε ∈ , t s> : J. Foukzon et al. DOI: 10.4236/***.2019.***** 38 Journal of Modern Physics ( ) ( ), ,, , .t q t s s qH F H Fε δ ε−≤ (195) Proof. Consider any δ-cover { }i iU ∈ of F. Then each t s t s iU δ − −≤ since iU δ≤ , so: .t t s s st si i i iU U U Uδ − −= ≤ (196) From (196) it follows that t st s i i q q i i U U x x δ ε ε − ≤ + + (197) and summing (196) over all i∈ we obtain 1 1 . t s i it s q qi i i i U U x x δ ε ε ∞ ∞− = = ≤ + + ∑ ∑ (198) Thus (195) follows by taking the infimum. Theorem 3.4.2. 1) If ( )( ), ,s qH F ε ε < ∞     , and if t s> , then ( )( ), , 0t qH F ε ε =   . 2) If ( )( ),0 ,s qH F ε ε<     , and if t s< , then ( )( ), ,t qH F ε ε = ∞   . Proof. 1) The result follows from (195) after taking limits, since ( ]0,1ε∀ ∈ by definitions follows that ( ), ,s qH F ε < ∞ . 2) From (3.4.10) ( ]0,1 , 0ε δ∀ ∈ ∀ > follows that ( ) ( ), ,1 , , .s q t qs t H F H Fε εδ − ≤ (199) After taking limit 0δ → , we obtain ( ), ,t qH F ε = ∞ , since ( ) 10lim s tδ δ −− → = ∞ and ( ) ( ), ,0lim , , 0s q s qH F H Fδ δ ε ε→ = > . Definition 3.4.6. We define now the Hausdorff-Colombeau dimension ( )dim ,HC F q of a set F (relative to 0q > ) as ( ) ( ) ( )( ){ } ( ) ( )( ){ } , , dim , sup | , inf = | , 0 . s q HC s q F q s s q H F s s q H F ε ε ε ε = = = ∞ = =     (200) Remark 3.4.1. From theorem 3.4.2 follows that for any fixed q q=  : ( )( ), , 0s qH F ε ε =    or ( )( ), ,s qH F ε ε = ∞    everywhere except at a unique value s, where this value may be finite. As a function of s, ( ), ,s qH F ε  is decreasing function. Therefore, the graph of ( ), ,s qH F ε  will have a unique value where it jumps from ∞ to 0. Remark 3.4.2. Note that the graph of ( )( ), ,s qH F ε ε  for a fixed q q=  is ( )( ) ( ) ( ) ( ) , if dim , , 0 if dim , undetermined if dim , HC s q HC HC s F q H F s F q s F q ε ε ∞ > = >  =         (201) Definition 3.4.7. We say that fractal nF has a negative dimension relaJ. Foukzon et al. DOI: 10.4236/***.2019.***** 39 Journal of Modern Physics tive to 0q > if ( )dim , 0HC F q q− < . 4. Scalar Quantum Field Theory in Spacetime with Hausdorff-Colombeau Negative Dimensions 4.1. Equation of motion and Hamiltonian Scalar quantum field theory and quantum gravity in spacetime with noninteger positive Hausdorff dimensions developed in papers [28] [29] [30] [31]. Quantum mechanics in negative dimensions developed in papers [32] [33] Scalar quantum field theory and quantum gravity in spacetime with Hausdorff-Colombeau negative dimensions originally developed in paper [12]. In this section only free scalar quantum field in spacetime with negative dimensions briefly is considered. A negative-dimensional spacetime structure is a desirable feature of superrenormalizable spacetime models of quantum gravity, and the most simply way to obtain it is to let the effective dimensionality of the multifractal universe to change at different scales. A simple realization of this feature is via suitable extended fractional calculus and the definition of a fractional action. Note that below we use canonical isotropic scaling such that: 1, 0,1, , 1x Dμ μ  = − = −  t (202) while replacing the standard measure with a nontrivial Colombeau-Stieltjes measure, ( )( ) [ ] [ ) t f t d d d , , , 1, . D Dx x x D ε η ε η α α → = = ⋅ ∈ −∞ (203) Here tD is the topological (positive integer) dimension of embedding spacetime and α is a parameter. Any Colombeau integrals on net multifractals can be approximated by the left-sided Colombeau-Riemann--Liouville complex milti-fractional integral of a function ( )tL : ( ) ( )( ) ( ){ }( ) ( ) ( ) ( )( ) ( ) ( )( ) ( ) ( ) ( ) ( )( ) 1 , 10 d , d , , , 1 i i ii z t t tmz t t i i z tz t i t t i x t I t t z t t t t i t z t ε εεε ε ε ε ε η ε ε η ε − =  − +   ∝  Γ    − − +   =  Γ +   ∑∫ ∫L L (204) where ( ]0,1ε ∈ , t is fixed and the order ( )z t is (related to) the complex Hausdorff-Colombeau dimensions of the set. In particular if , 1, 2, ,iz i m∈ =  is a complex parameter an integral on net multifractals can be approximated by finite sum of the left-sided Colombeau-Riemann-Liouville complex fractional integral of a function ( )tL J. Foukzon et al. DOI: 10.4236/***.2019.***** 40 Journal of Modern Physics ( ) ( )( ) { } ( ) ( ) ( ) ( ) ( )( ) ( )( ) 1 ,0 1 ,1 1 1 d , 1 d , , . 1 m i i ii ii t z t t zm mz ti i i zz m i i x t I I t t i t z t t t i t z ε ε ε ε εε ε ε ε η ε ε ε η ε = − = = =  ∝       = − +    Γ   − − +   =  Γ +   ∫ ∑ ∑ ∫ ∑  L L (205) Note that a change of variables t t t→ − transforms Equation (205) into the form ( ) ( )( ) [ ]( )( ) ( ) 1 10 0 d , d . iz t tm i t i x t t t t z tε ε ε η ε − =  +  = −  Γ  ∑∫ ∫L L (206) The Colombeau-Riemann-Liouville multifractional integral (206) can be mapped onto a Colombeau-Weyl multifractional integral in the formal limit t → +∞ . We assume otherwise, so that there exists ( )lim t z t→+∞ and ( ) ( ) ( )lim ,t t t q t q t→+∞ − =   L L . In particular if z∈ is a complex parameter a change of variables t t t→ − transforms Equation (206) into the form ( ) [ ]( ) ( ) ( ) 1 ,1 1 d , . i i z tm mz ti i i t i I t q t q t zε εε ε ε − = =  +  =    Γ  ∑ ∑ ∫ L (207) This form will be the most convenient for defining a Colombeau-Stieltjes field theory action. In tD dimensions, we consider now the action ( ) ( ) ( ) ( )( )d , , ,M x x xε ε μ εε εη ε φ φ = ∂ ∫S L (208) where , μφ φ ∂ L is the Lagrangian density of the scalar field ( )( )xε εφ and where ( )( ) ( )( ) ( )( )t 1 , ,1 0d , , d , , : , Dm i iix f x x f x M μ μ μμε ε ε η ε ε ε− = = = →∑ ∏  (209) is some Colombeau-Stieltjes measure. We denote with pair ( )( )( ), d ,M x εη ε the metric spacetime M equipped with Colombeau-Stieltjes measure ( )( )d ,x εη ε . The former can be taken to be the canonical scalar field Lagrangian, ( ) ( )( ) ( ) ( )( )1, ,2x x V μ ε μ ε μ ε ε ε εεε φ φ φ φ φ ∂ = − ∂ ∂ − L (210) where ( )V φ is a potential and contraction of Lorentz indices is done via the Minkowski metric ( ), , ,μν μνη = − + + . As for the Colombeau-Stieltjes measure, we make the multifractal spacetime isotropic choice ( )( ) ( ), t, , , 1, , 1; 1, , .iif f D i mεμ ε εε μ= = − =  (211) Hence the scalar field action (208) reads ( ) ( ) ( ) ( )( ) ( ) ( )t ,1 d , , 1d , 2 M m D jj x x x xv x V ε ε μ εε ε μ ε μ ε ε ε ε η ε φ φ φ φ φ =  = ∂    = ∂ ∂ +     ∫ ∑ ∫ S (212) where ( )( )v xε ε is a coordinate-dependent J. Foukzon et al. DOI: 10.4236/***.2019.***** 41 Journal of Modern Physics Lorentz scalar ( )( ) ( ) ( )t , 1 1 .j D j v x s x ε αε ε ε −    =     +   (213) We define now the Dirac distribution as Colombeau generalized function by equation ( ) { } ( ) ( )( )f1 d , , .jjDm jj vx x mεη ε δ ε= =∑ ∫ , (214) In particular for the case 1m = ( ) { } ( ) ( )( )fd , , 1.Dvx x εη ε δ ε =∫ (215) Invariance of the action under the infinitesimal shift ( ) ( ) ( )x x xφ φ δφ→ + gives the equation of motion for a generic weight ( ), , 1, ,iv i mε ε =  : ( ) , 1 , d 0. d m i i i v v x μ ε μ ε ε μ εε ε φ φ=    ∂ ∂ ∂ − + =      ∂ ∂ ∂       ∑L L (216) In particular for the case 1m = we obtain ( ) d 0. d v v x μ ε μ ε ε μ εε ε φ φ  ∂   ∂ ∂ − + =    ∂ ∂ ∂     L L (217) From Equation (212) and Equation (216) we obtain ( ) ( ),1 , d 0. d m i i i v V v μ ε μ ε ε εε ε ε εε φ φ φ φ=    ∂   + ∂ − =              ∑ (218) where μμ= ∂ ∂ . In particular for the case 1m = we obtain ( ) ( )d 0. d v V v μ ε μ ε ε εε ε ε εε φ φ φ φ  ∂    + ∂ − =           (219) 4.2. Propagator in Configuration Space with Negative-Dimensions We define the canonical vacuum-to-vacuum amplitude by [ ]( ) ( )( ),1, exp d ,m jjZ J i Jε ε εε εε φ η φ= = + ∑∫ ∫D L (220) where J is a source. Integration by parts in the exponent leads to the Lagrangian density for a free field as ( ) ( ), 21 , 1 1 , 2 2 m j j j v m v μ ε μ ε ε ε ε ε εε ε ε ε φ φ φ φ =   ∂  = + ∂ − = I      ∑L (221) where , 2 1 , ; 1, , .m jj j v m j m v μ ε μ ε ε = ∂ I = + ∂ − =∑  (222) In particular for the case 1m = we obtain J. Foukzon et al. DOI: 10.4236/***.2019.***** 42 Journal of Modern Physics 2 . v m v μ ε μ ε ε ∂ I = + ∂ − (223) The propagator is the Green function ( )( )G xε ε solving the equation ( )( ) ( )( ), ,DvG x xε ε ε εδ ε − I = (224) where ( )t 1 0D D α− = − < . By virtue of Lorentz covariance, the Green function ( )G xε must depend only on the Lorentz interval 2 2iis x x x x tμμ= = − , where 0x t= and t1, , 1i D= − . In particular, ( ) ( )( )( )v v s xε εε ε= with the correct scaling property is ( )( )( ) ( ) ( ) 1 , . D v s x s x s x x xμε με ε ε − −  = + =      (225) Note that ( ) ( ) 2 t 1, .s s s x D s s μ μ ε εε ε − ∂ = ∂ = ∂ + ∂ + +  (226) Hence the inhomogeneous Equation (224) reads ( ) ( )( ) ( )( ) 2 2t 1 , .Ds s v D m G x x s ε ε ε ε α δ ε ε − − ∂ + ∂ − =  +  (227) We first consider the Euclidean propagator and denote with 2iir x x t= + the Wick-rotated Lorentz invariant. In the massless case, the solution of the homogeneous equations for any 0α < is ( )( ) t2 2, . 2 D G r Cr βε ε α β + = = (228) Let us now consider the massive case.The solution of the homogeneous equation ( )( ) 0G rε ε εI = for any 0α < is ( )( ) ( ) ( ) 2 t 2 t t1 22 2 2 2 , D D D rG r C mr C mr m α ε α αε + ⋅ + ⋅ + ⋅ − −   = +        K I (229) where 1 2,C C are constants and λK and λI are the modified Bessel functions. The function ( )zνI is ( ) ( )( ) 2 0 2 . ! 1 k k z z k k ν ν ν + ∞ = = Γ + +∑I (230) Formula (230) is valid providing 1, 2, 3,ν ≠ − − −  ( ) ( )( ) 2 0 2 ! 1 k k z z k k ν ν ν − + ∞ − = = Γ − + + ∑I (231) Formula (231) is obtained by replacing ν in (232) with a ν− . ( ) ( ) ( )π . 2sin π z z zν ν νν− −  = − − K I I (232) The modified Bessel functions ( )zν−I and J. Foukzon et al. DOI: 10.4236/***.2019.***** 43 Journal of Modern Physics ( )zν−K have the following asymptotic forms for 0z → : ( ) ( ) ( ) ( ) 1 2 1, , 1, 2, 3, 2 21 zz z z ν ν ν νν νν − − − −    Γ − ≠ − − −   Γ − +    K I  (233) Since for small 0m the solution must agree with the massless case (228), we can set 2 0C = . To find the solution of the inhomogeneous equation, one exploits the fact that the mass term does not contribute near the origin. Expanding Equation (229) at 0mr  when 0α < ( 2 0C = ), we find ( )( ) ( ) t t4 2 t 22 2 1 2 2 2 i iD D iDG r C r α α ε ε α+ ⋅ + ⋅−  + ⋅ = Γ −    (234) which must coincide with Equation (228). This gives the coefficient 1C and the propagator reads ( ) ( ) t tt t 2 2 2 t 22 1 2 . 2 2π 2 i iD DD D mG r mr rD α αα + ⋅ − + −  Γ    = −      Γ −    , K (235) 5. The Solution Cosmological Constant Problem 5.1. Einstein-Gliner-Zel'dovich Vacuum with Tiny Lorentz Invariance Violation We assume now that: 1) Poincaré group of momentum space is deformed at some fundamental high-energy cutoff ∗Λ [9] [10]. 2) The canonical quadratic invariant 2 ab a bp p pη= collapses at high-energy cutoff ∗Λ and being replaced by the non-quadratic invariant: ( ) 2 0 . 1 ab a bp pp l p η ∗Λ = + (236) 3) The canonical concept of Minkowski space-time collapses at a small distance 1l ∗ − Λ ∗= Λ to fractal space-time with Hausdorff-Colombeau negative dimension and therefore the canonical Lebesgue measure 4d x being replaced by the Colombeau-Stieltjes measure ( )( ) ( )( )( )4d , d ,x v s x xεε εη ε = (237) where ( )( )( ) ( ) ( ) 1 , , D v s x s x s x x x ε ε ε μ μ ε − −  = +      = (238) see subsection IV.2. 4) The canonical concept of momentum space collapses at fundamental J. Foukzon et al. DOI: 10.4236/***.2019.***** 44 Journal of Modern Physics high-energy cutoff ∗Λ to fractal momentum space with Hausdorff-Colombeau negative dimension and therefore the canonical Lebesgue measure 3d k , where ( ), ,x y zk k k=k being replaced by the Hausdorff-Colombeau measure ( ) ( ) ( ) 1, d d= , D D D D D D D D D p p d p ε ε ε ε + + + − − − − + − −∆ ∆ ∆    + +        k k k (239) where ( ) ( ) 22π 2 D D D ± ± ± ∆ = Γ and x y zp k k k= = + +k . Remark 5.1.1. Note that the integral over measure ,dD D + − k is given by formula (185). Thus vacuum energy density ( )eff, , ,D D pε μ+ − ∗ for free quantum fields is ( ) ( ) ( ) ( )eff eff eff eff, , , , , , , .D D p p D D pε μ ε μ ε μ ε μ+ − + −∗ ∗ ∗= + +  (240) Here the quantity ( )effε μ is given by formula ( ) ( ) ( ) ( ) ( ) eff eff eff 2 2 3 eff 3 0 2 2 2 0 2 2 2 0 0 1 d d 2 2π d d d d p f K f p p p K f p p p μ μ μ μ μ μ ε μ μ μ μ μ μ μ μ μ μ ≤ ≤ = + = + = + ∫ ∫ ∫ ∫ ∫ ∫  k k k (241) where ( )3 2π , 1 2π K c= =  . The quantity ( )eff , pε μ ∗ is given by formula ( ) ( ) ( ) ( ) eff eff 2 2 3 eff 3 0 2 2 2 0 1, d d 2 2π d d . p p p f K f p p p μ μ μ μ ε μ μ μ μ μ μ μ ∗ ∗ ∗ < < < < = + = + ∫ ∫ ∫ ∫  k k k k (242) The quantity ( )eff, , ,D D pε μ+ − ∗ (since Equation (22) holds) is given by formula ( ) ( ) ( ) eff eff 2 4 2 2 2 , 2 2 2 22 2 0 , , , 1d d , 1 11 D D p D D p l l K f l ll μ ε μ μ μ μ μ μ μ μμ + − ∗ ∗ ∗ ∗ ∗∗ + − ∗ Λ Λ ≥ Λ ΛΛ    = + + +  − −−  ∫ ∫  k k k (243) where ( )3 1 , 1 2 2π K c′ = =  . Remark 5.1.2. We assume now that 2 2 4 21, 1l lμ μ ∗ ∗Λ Λ   and therefore from Equation (243) we obtain J. Foukzon et al. DOI: 10.4236/***.2019.***** 45 Journal of Modern Physics ( ) ( ) ( ) eff eff 3,2 2 2 , 0 0 , , , d d d d .D D D p p D D p K l f K f μ μ ε μ μ μ μ μ μ μ + − − ∗ ∗ + − ∗ Λ ≥ ≥ ′ ′= + +∫ ∫ ∫ ∫ eff k k k k k (244) From Equation (244) and Equation (239) we obtain ( ) ( ) ( ) ( ) ( ) ( )( ) ( ) ( ) ( ) ( ) ( ) ( )( ) eff eff eff eff eff eff 2 , 2 2 , 0 0 1 2 0 2 2 1 0 2 1 0 , , , d d d d dd d d d d D D D D p p D p D D p D D D p D D p K l f K f p pK l D D f p p p p K D D f p K l D D f p p μ μ μ ε μ ε μ ε μ μ μ μ μ μ μ μ μ μ ε μ μ μ ε μ μ μ + − + − ∗ ∗ + −∗ + −∗ − + + − ∗ Λ ≥ ≥ − ∞+ − Λ − ∞+ − + − + − Λ ′ ′= + + ′= ∆ ∆  +    + ′+ ∆ ∆  +    ′= ∆ ∆ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ k k k k k ( ) ( ) ( )eff 2 2 10 d d .D DpK D D f p p p μ μ μ μ ∗ − + ∗ ∞ ∞+ − + −′+ ∆ ∆ + ∫ ∫ ∫ (245) Remark 5.1.2. We assume now that: 2 6.D D− ++ + ≤ − (246) Note that ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff eff e 2 2 1 0 2 20 2 1 0 0 4 3 4 0 1 2 0 0 d d d 1 d 1d d d d 2 1 d d 8 d d 1 2 D D p D D p D D D D p p D D D D p D D D D f p p p f p p p f p p f p p f p p O p p pf f D D D D μ μ μ μ μ μ μ μ μ μ μμ μ μ μ μ μ μ μ μ μ μ μ μ μ μ − + ∗ − + ∗ − + − + ∗ ∗ − + − + ∗ − + − + ∞ + − ∞ + ∞ ∞+ + − ∞ + − + − ∗ + + + ∗ ∗ − + − + + = + = + − + = + + + + ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ( ) ( ) ( ) ff eff 1 4 4 0 d . 8 1 D D D Dp f O p D D μ μ μ μ − + − + + − + −∗ ∗− + − + + − ∫ ∫ (247) Thus finally we obtain ( ) ( ) ( ) ( ) ( )( ) ( ) ( ) ( ) eff eff eff eff 1 0 2 0 2 4 4 0 , , , d 1 0.5 d d . 8 1 D D D D D D D D D D p K p f D D pK l D D f D D K p f O p D D μ μ μ ε μ μ μ μ μ μ μ μ μ − + − + − + − + + − ∗ + + ∗ − + + + − ∗ Λ − + + − + −∗ ∗− + ′ = + +  ′+ ∆ ∆ +  + ′ − + + − ∫ ∫ ∫ (248) J. Foukzon et al. DOI: 10.4236/***.2019.***** 46 Journal of Modern Physics Remark 5.1.3. Note that (see Equations (42)): ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff eff eff eff eff 4 2 2 4 1 0 0 0 4 6 8 5 2 0 0 0 , , 1 1 1d d ln d 4 4 8 1 1 1ln d d . 8 32 p p p f p f C p f f f O f p p μ μ μ μ μ μ ε μ ε μ ε μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ ∗ ∗ ∗ ∗ ∗ − ∗ ∗ = +  = + + −       + − +         ∫ ∫ ∫ ∫ ∫ ∫  (249) From Equation (240), Equation (248) and Equation (249) finally we obtain ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff eff eff eff eff eff 4 2 2 4 1 0 0 0 4 6 8 5 2 0 0 0 2 , , , , , , , 1 1 1d d ln d 4 4 8 1 1 1ln d d 8 32 .D D D D p p D D p p f p f C p f f f O f p p O p μ μ μ μ μ μ ε μ ε μ ε μ ε μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ + − + − ∗ ∗ ∗ ∗ ∗ ∗ − ∗ ∗ − ++ + ∗ = + +  = + + −       + − +         + ∫ ∫ ∫ ∫ ∫ ∫  (250) The pressure ( )eff, , ,p D D pμ+ − ∗ for free scalar quantum field is ( ) ( ) ( ) ( )eff eff eff eff, , , , , , , .p D D p p p p p D D pμ μ μ μ+ − + −∗ ∗ ∗= + +  (251) Here the quantity ( )effp μ is given by formula ( ) ( )eff 4 eff 0 2 2 d d . 3 p K pp f p p μ μ μ μ μ μ< = + ∫ ∫ (252) The quantity ( )eff ,p pμ ∗ is given by formula ( ) ( )eff 4 eff 0 2 2 , d d . 3 p p K pp p f p p μ μ μ μ μ μ∗ ∗ ≤ ≤ = + ∫ ∫ (253) The quantity ( )eff, , ,p D D pμ+ − ∗ is given by formula ( ) ( )eff 4 eff 0 2 2 , , , d d , 3 p p K pp D D p f p p μ μ μ μ μ∗ + − ∗ > ′ + ∫ ∫   (254) where ( )3 1 , 1 2 2π K c′ = =  . Remark 5.1.4. Note that (see Equations (42): ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff eff eff eff eff 4 2 2 0 0 4 4 2 0 0 6 8 5 2 0 0 , , 1 1d d 12 12 1 1ln d ln d 8 8 5 1 d . 32 p p p p p p f p f C p f f f O f p p μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ ∗ ∗ ∗ ∗ ∗ − ∗ ∗ = + = −  + + −       + +         ∫ ∫ ∫ ∫ ∫ ∫  (255) From Equation (250), Equation (254) and Equation (255) similarly as above finally we get J. Foukzon et al. DOI: 10.4236/***.2019.***** 47 Journal of Modern Physics ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff eff eff 4 2 2 0 0 4 4 2 0 0 6 8 5 2 2 0 0 , , , 1 1d d 12 12 1 1ln d ln d 8 8 5 1 d . 32 D D p D D p p f p f C p f f f O f p O p p μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ − + + − ∗ ∗ ∗ ∗ − + + ∗ ∗ ∗ = −  + + −       + + +        ∫ ∫ ∫ ∫ ∫ ∫ (256) Remark 5.1.5. We assume now that: ( ) ( ) ( ) eff eff eff 2 4 0 0 0 d d d 0.f f f μ μ μ μ μ μ μ μ μ μ μ= = =∫ ∫ ∫ (257) From Equation (250), Equation (256) and Equation (257) finally we get ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff 4 2 eff 0 4 2 eff 0 1, , , ln d , 8 1, , , ln d . 8 D D p f O p p D D p f O p μ μ ε ε μ μ μ μ μ μ μ μ μ μ + − − ∗ ∗ + − − ∗ ∗ = + = − + ∫ ∫   (258) Remark 5.1.6. Note that the Equation (258) can be obtained without fine-tuning (257) which was assumed in Zel'dovich paper [1]. In order to obtain Equation (5.1.23) and strictly weaker conditions we assume now that: 1) ( ) ( ) ( ). . . . eff ,ns m g mf f fμ μ μ μ−= + = (259) where 0n > is an parameter, ( ). .s mf μ corresponds to standard matter and where ( ). .g mf μ corresponds to physical ghost matter, see Equation (32). 2) ( ) ( ) ( ) eff eff eff 4 1 0 2 2 2 0 4 3 0 d 0, d 0, ln d 0 I p f I p f I p f μ μ μ μ μ μ μ μ μ μ μ ∗ ∗ ∗ = ≈ = ≈ = ≈ ∫ ∫ ∫ (260) 3) ( ) ( ) eff 4 1 2 3 0 ln d .I I I f μ μ μ μ μ+ + ∫ (261) 5.2. Zeropoint Energy Density Corresponding to a Non-Singular Gliner Cosmology We assume now that ( ) ( ) ( ) eff eff eff 4 2 eff 0 0 0 d 0, d 0, d 0, .f f f p μ μ μ μ μ μ μ μ μ μ μ μ∗= < >∫ ∫ ∫  (262) J. Foukzon et al. DOI: 10.4236/***.2019.***** 48 Journal of Modern Physics From Equation (250), Equation (256) and (262) we obtain ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff eff 2 2 4 1 0 0 4 6 2 0 0 8 5 2 0 , , , 1 1d ln d 4 8 1 1 1ln d d 8 32 ,D D D D p p f C p f f f p O f p O p μ μ μ μ μ ε ε μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ − + + − ∗ ∗ ∗ ∗ − + + ∗ ∗  = − −      + −       + +     ∫ ∫ ∫ ∫ ∫  (263) and ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff 2 2 4 2 0 0 4 6 2 0 0 8 5 2 0 , , , 1 1d ln d 12 8 1 5 1ln d d 8 32 D D p p D D p p f C p f f f p O f p O p μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ − + + − ∗ ∗ ∗ ∗ − + + ∗ ∗  = − − +      − +       + +     ∫ ∫ ∫ ∫ ∫  eff (364) correspondingly. From Equation (263) and Equation (264) we obtain ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff eff eff eff eff eff eff 2 2 4 2 0 0 4 6 2 0 0 2 2 4 1 0 0 4 6 2 0 0 4 1 33 d 3 ln d 4 8 3 5 3ln d d 8 32 1 1d ln d 4 8 1 1 1ln d d 8 32 1 ln d 4 p p f C p f f f p p f C p f f f p p f μ μ μ μ μ μ μ μ ε μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ ∗ ∗ ∗ ∗ ∗ ∗ ∗  + = − − +      − +      + − −      + −     = − ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ ( ) ( ) ( ) ( ) ( ) eff eff eff eff 4 2 1 0 0 4 6 2 0 0 3 d 1 5 1ln d d 0. 4 16 C C f f f p μ μ μ μ μ μ μ μ μ μ μ μ μ μ ∗ − +   − + <    ∫ ∫ ∫ ∫ (265) Therefore under conditions (262) the inequality 2 3 0pε ε− < + < (266) corresponding to Gliner non-singular cosmology [2] [4] is satisfied. 5.3. Zeropoint Energy Density in Models with Supermassive Physical Ghost Fields We assume now that: 1) ghost fields corresponding to massive spin-2 particle with mass 2m and to massive scalar particle with mass 0m appears (see Subsection II.2) as real physJ. Foukzon et al. DOI: 10.4236/***.2019.***** 49 Journal of Modern Physics ical fields in action () Remark 5.3.1. Note that their unphysical behavior may be restricted to arbitrarily high-energy cutoff Λ by an appropriate limitation on the renormalized masses 2m and 0m . Actually, it is only the massive spin-two excitations of the field which give the problem with unitarity and thus require a very large mass (see Subsection II.2). 2) Poincaré group is deformed at some fundamental high-energy cutoff ∗Λ ( ) 2 20 2 0 2, .m m m c m c∗ ∗Λ = Λ < (267) The canonical quadratic invariant 2 ab a bp p pη= collapses at high-energy cutoff ∗Λ and being replaced by the non-quadratic invariant ( ) 2 0 . 1 ab a bp pp l p η ∗Λ = + (268) 3) The canonical concept of Minkowski space-time collapses at a small distance to fractal space-time with Hausdorff-Colombeau negative dimension and therefore the canonical Lebesgue measure 4d x being replaced by the Colombeau-Stieltjes measure ( )( ) ( )( )( )4d , d ,x v s x xεε εη ε = (269) where ( )( )( ) ( ) ( ) 1 , , D v s x s x s x x xμε με ε ε − −  = + =      (270) 4) we assume that ( ) ( ) ( ). . . . ,s m g mf f fμ μ μ= + (271) where ( ). .s mf μ corresponds to standard matter and where ( ). .g mf μ corresponds to physical ghost matter. Remark 5.3.2. We assume now that ( ) ( ) 1 2 0 eff eff 2 2 eff , 1 0 nO n m c m c f μ μ μ μ μ μ μ − > ≤ ≤=  >   (272) Thus vacuum energy density ( )1 2eff eff, , ,D Dε μ μ+ − for free quantum fields is ( ) ( ) ( )1 2 1 2 1 2eff eff eff eff eff eff, , , , , , , .D D D Dε μ μ ε μ μ ε μ μ+ − + −= +  (273) Here the quantity ( )1 2eff eff,ε μ μ is given by formula ( ) ( ) ( ) ( ) 2 eff 1 eff 2 eff 1 eff 1 2 2 2 3 eff 3 2 2 2 1, d d 2 2π d d , p f K f p p p μ μ μ μ μ μ ε μ μ μ μ μ μ μ μ ≤ ≤ = + = + ∫ ∫ ∫ ∫  eff k k k (274) where ( )3 2π , 1 2π K c= =  . The quantity ( )1 2eff eff, , ,D Dε μ μ+ − is given by formula J. Foukzon et al. DOI: 10.4236/***.2019.***** 50 Journal of Modern Physics ( ) ( ) ( )2eff1 eff 1 2 eff eff 4 22 2 2 , 2 2 2 22 2 , , , 1d d , 1 11 D D D D llK f l ll μ μ μ ε μ μ μμ μ μ μ μ μμ + − ∗ ∗∗ + − ΛΛ Λ Λ> Λ    ′= + + +  − −−  ∫ ∫ k k k  (275) where ( )3 1 , 1 2 2π K c′ = =  . Remark 5.3.2. We assume now that 2 2 1lμ Λ∗ < , and therefore from Equation (5.3.9) we obtain ( ) ( ) ( ) 2 2 eff eff 1 1 eff eff 1 2 eff eff 3,2 2 2 , , , , d d d d .D D D D D K l f K f μ μ μ μ μ μ ε μ μ μ μ μ μ μ μ + − − + − Λ > > ′ ′+ +∫ ∫ ∫ ∫ k k k k k (276) From Equation (276) and Equation (239) we obtain ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 2 2 eff eff 1 1 eff eff 2 eff 1 eff 2 eff 1 eff 1 2 eff eff 2 , 2 2 , 1 2 2 2 1 , , , d d d d dd d d D D D D D ' D D D D D K l f K f p pK D D l f p p p p K D D f p μ μ μ μ μ μ μ μ μ ε μ μ μ ε ε μ μ μ μ μ μ μ μ μ μ μ ε μ μ μ ε + − + − + − + − + − Λ > > − ∞+ − Λ − ∞+ − ′ ′+ +      = ∆ ∆   +      + ′+ ∆ ∆  +    ∫ ∫ ∫ ∫ ∫ ∫ ∫ ∫ k k k k k ( ) ( ) ( ) ( ) ( ) ( ) 2 eff 1 eff 2 eff 1 eff 2 1 2 2 1 d d d d . D D D D K D D l f p p K D D f p p p μ μ μ μ μ μ μ μ μ μ μ μ − + − + ∞+ − + − Λ ∞+ − + −            ′= ∆ ∆     ′+ ∆ ∆ +   ∫ ∫ ∫ ∫ (277) Note that 2 2 4 6 2 2 2 2 4 6 2 4 6 3 5 1 1 11 1 2 8 16 1 1 1 2 8 16 p p p pp p p p μ μ μ μ μ μ μ μ μ μ μ   + = + = + − + +    = + − + +   (278) By inserting Equation (278) into Equation (274) we get ( ) ( ) ( ) ( ) 2 eff 1 eff 2 eff 1 eff 2 eff 1 eff 1 2 eff eff 2 4 6 2 3 5 4 6 8 2 3 5 0 3 5 7 9 3 5 0 , 1 1 1d d 2 8 16 1 1 1d d 2 8 16 1 1 1d 3 2 5 8 167 9 p p p pK f p p p p pK f p p p p p pK f μ μ μ μ μ μ μ μ μ ε μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ ≤   = + − + +       = + − + +         = + − + +    ∫ ∫ ∫ ∫ ∫    J. Foukzon et al. DOI: 10.4236/***.2019.***** 51 Journal of Modern Physics ( ) ( ) ( ) ( )( ) 2 eff 1 eff 2 eff 1 eff 2 eff 1 eff 3 5 1 9 2 2 2 2 3 5 5 3 1 1 2 2 2 2 5 3 1 1 1 212 2 2 2 eff 1 1 1d 3 2 5 8 167 9 1 1 1 1d 3 10 56 144 1 1 1 1d . 3 10 56 144 n K f K f K f o μ μ μ μ μ μ μ μ μ μμ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ − − − +    = + − + +       = + − + +      = + − + +    ∫ ∫ ∫   (279) The pressure ( )1 2eff eff, , ,p D D μ μ+ − for free quantum fields is ( ) ( ) ( )1 2 1 2 1 2eff eff eff eff eff eff, , , , , , , .p D D p p D Dμ μ μ μ μ μ+ − + −= +  (280) Here the quantity ( )1 2eff eff,p μ μ is given by formula ( ) ( ) ( ) ( ) 2 eff 1 eff 2 eff 1 eff 2 1 2 3 eff eff 3 2 2 4 2 2 1, d d 2 2π d d . 3 p p f K pf p p μ μ μ μ μ μ μ μ μ μ μ μ μ μ ≤ ≤ = + = + ∫ ∫ ∫ ∫ k k k k (281) The quantity ( )1 2eff eff, , ,p D D μ μ+ − is given by formula ( ) ( ) 2 eff 1 eff 2 1 2 , eff eff 2 2 , , , d d , 3 D D p Kp D D f μ μ μ μ μ μ μ μ + −+ − > ′ + ∫ ∫  k k k  (282) where ( )3 1 , 1 2 2π K c′ = =  . Note that 1 2 1 22 2 2 4 6 1 2 4 6 2 4 6 3 5 7 1 1 1 3 51 2 8 16 1 1 3 5 2 8 16 p p p p p p p p μ μμ μ μ μ μ μ μ μ μ − − −    = +  +     = − + − +    = − + − +   (283) By inserting Equation (283) into Equation (281) we get ( ) ( ) ( ) ( ) 2 eff 1 eff 2 eff 1 eff 2 eff 1 eff 1 2 eff eff 2 4 6 4 3 5 7 4 6 8 10 3 5 7 5 7 9 11 3 5 7 0 , 1 1 3 5d d 3 2 8 16 1 3 5d d 3 2 8 16 1 3 5d 3 5 2 8 167 9 10 p p p K p p pf p p K p p p pf p K p p p pf μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ ≤ ≤   = − + − +      = − + − +      = − + − +    ∫ ∫ ∫ ∫ ∫    J. Foukzon et al. DOI: 10.4236/***.2019.***** 52 Journal of Modern Physics ( ) ( ) ( ) ( )( ) 2 eff 1 eff 2 eff 1 eff 2 eff 1 eff 5 7 9 11 2 2 2 2 3 5 7 3 1 1 3 2 2 2 2 3 1 1 3 1 212 2 2 2 eff 1 3 5d 3 5 2 8 167 9 10 1 1 1 1d 3 5 14 24 32 1 1 1 1d . 3 5 14 24 32 n K f K f K f o μ μ μ μ μ μ μ μ μ μμ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ − − − − − −    = − + − +       = − + − +      = − + − +    ∫ ∫ ∫   (284) 6. Discussion and Conclusion We will now briefly review the canonical assumptions that are made in the usual formulation of the cosmological constant problem. The canonical assumptions: 1) The physical dark matter. Dark matter is a hypothetical form of matter that is thought to account for approximately 85% of the matter in the universe, and about a quarter of its total energy density. The majority of dark matter is thought to be non-baryonic in nature, possibly being composed of some as-yet-undiscovered subatomic particles. Its presence is implied in a variety of astrophysical observations, including gravitational effects that cannot be explained unless more matter is present than can be seen. For this reason, most experts think dark matter to be ubiquitous in the universe and to have had a strong influence on its structure and evolution. The name dark matter refers to the fact that it does not appear to interact with observable electromagnetic radiation, such as light, and is thus invisible (or 'dark') to the entire electromagnetic spectrum, making it extremely difficult to detect using usual astronomical equipment. Because dark matter has not yet been observed directly, it must barely interact with ordinary baryonic matter and radiation. The primary candidate for dark matter is some new kind of elementary particle that has not yet been discovered, in particular, weakly-interacting massive particles (WIMPs), or gravitationally-interacting massive particles (GIMPs). Many experiments to directly detect and study dark matter particles are being actively undertaken, but none has yet succeeded. 2) The total effective cosmological constant effλ is on at least the order of magnitude of the vacuum energy density generated by zero-point fluctuations of the standard particle fields. 3) Canonical QFT is an effective field theory description of a more fundamental theory, which becomes significant at some high-energy scale ∗Λ . 4) The vacuum energy-momentum tensor is Lorentz invariant. 5) The Moller-Rosenfeld approach [34] [35] to semiclassical gravity by using an expectation value for the energy-momentum tensor is sound. 6) The Einstein equations for the homogeneous Friedmann-Robertson-Walker metric accurately describes the large-scale evolution of the Universe. Remark 6.1.1. Note that obviously there is a strong inconsistency between Assumptions 2 and 3: the vacuum state cannot be Lorentz invariant if modes are J. Foukzon et al. DOI: 10.4236/***.2019.***** 53 Journal of Modern Physics ignored above some high-energy cutoff ∗Λ , because a mode that is high energy in one reference frame will be low energy in another appropriately boosted frame. In this paper Assumption 3 is not used and this contradiction is avoided. Remark 6.1.2. Note that also, Assumptions 1, 3, 4 and 5 is modified, which we denote as Assumptions 4 and 5 respectively. Modified assumptions 1') The physical dark matter. 2') The total effective cosmological constant effλ is on at least the order 5 eff effln nμ μ− + of magnitude of the renormalized vacuum energy density generated by zero-point fluctuations of standard particle fields and ghost particle fields, see subsection 1.2. 3') The vacuum energy-momentum tensor is not Lorentz invariant. 6.1. The Physical Ghost Matter and Dark Matter Nature In the contemporary quantum field theory, a ghost field, or gauge ghost is an unphysical state in a gauge theory. Ghosts are necessary to keep gauge invariance in theories where the local fields exceed a number of physical degrees of freedom. For example in quantum electrodynamics, in order to maintain manifest Lorentz invariance, one uses a four-component vector potential ( )A xμ , whereas the photon has only two polarizations. Thus, one needs a suitable mechanism in order to get rid of the unphysical degrees of freedom. Introducing fictitious fields, the ghosts, is one way of achieving this goal. Faddeev-Popov ghosts are extraneous fields which are introduced to maintain the consistency of the path integral formulation. Faddeev-Popov ghosts are sometimes referred to as "good ghosts". "Bad ghosts" represent another, more general meaning of the word "ghost" in theoretical physics: states of negative norm, or fields with the wrong sign of the kinetic term, such as Pauli-Villars ghosts, whose existence allows the probabilities to be negative thus violating unitarity. (VI.1) In contrary with standard Assumption 1 in the case of the new approach introduced in this paper we assume that: (VI.1.1.a) The ghosts fields and ghosts particles with masses at a scale less then a fixed scale effm really exist in the universe and formed dark matter sector of the universe, in particular: (VI.1.1.b) these ghosts fields give additive contribution to a full zero-point fluctuation (i.e. also to effective cosmological constant effλ [5], see subsection 1.2). (VI.1.1.c) Pauli-Villars renormalization of zero-point fluctuations (see subsection 1.2) is no longer considered as an intermediate mathematical construct but obviously has rigorous physical meaning supported by assumption (I.a-b). (VI.1.2) The physical dark matter formed by ghosts particles; (VI.1.3) The standard model fields do not to couple directly to the ghost sector in the ultraviolet region of energy at a scale less then a fixed large energy scale ∗Λ , in particular: J. Foukzon et al. DOI: 10.4236/***.2019.***** 54 Journal of Modern Physics (VI.1.3.a) The "bad" ghosts fields with masses at a scale less then a fixed scale effm , where 2 effm c ∗Λ , cannot appear in any effective physical lagrangian which contains also the standard particles fields. In additional though not necessary we assume that: (VI.1.4) The "bad" ghosts fields with masses at a scale m∗ , where 2m c∗ ∗Λ can appear in any effective physical lagrangian which contains also the standard particles fields, in particular: (VI.1.4.a) Pauli-Villars finite renormalization with masses of ghosts fields at a scale m∗ of the S-matrix in QFT (see Subsection 2.1-2) is no longer considered as an intermediate mathematical construct but obviously has rigorous physical meaning supported by assumption (IV). (VI.1.4.b) If the "bad" ghosts fields coupled to matter directly, it gives rise to small and controllable violetion of the unitarity condition. Remark 6.1.3. We emphasize that in universe standard matter coupled with a physical ghost matter has the equation of state [3]: ( ) ( ) ( ) ( ) eff 4 4 vac vac eff eff 0 1 ln d , 8 8π c p f G μ λ ε μ μ μ μ μ μ= − = =∫ (285) where ( ) ( ) eff eff , 1 0 nO n f μ μ μ μ μ μ − > ≤=  > (286) and where eff effm cμ = (see subsection I.2, Equation (46)) and therefore gives rise to a de Sitter phase of the universe even if bare cosmological constant 0λ = . (VI.1.5) In order to obtain QFT description of the dark component of matter in natural way we expand the standard model of particle physics on a sector of ghost particles, see [12], Section 2.3.2. QFT in a ghost sector developed in [12], Section 3.1-3.4 and Section 4.1-4.8. 6.2. Different Contributions to effλ . The total effective cosmological constant effλ is on at least the order of magnitude of the vacuum energy density generated by zero-point fluctuations of standard particle fields. Assumption 2 is well justified in the case of the traditional approach, because the contribution from zero-point fluctuations is on the order of 1 in Planck units and no other known contributions are as large thus, assuming no significant cancellation of terms (e.g. fine tuning of the bare cosmological constant λ ), the total effλ should be at least on the order of the largest contribution [15]. (VI.2) In contrary with standard Assumption 1 in the case of the new approach introduced in this paper we assume that: (VI.2.1) For simplicity though not necessary bare cosmological constant 0λ = . (VI.2.2) The total effective cosmological constant effλ depends only on mass distribution ( )f μ and constant eff effm cμ = but cannot depend on large J. Foukzon et al. DOI: 10.4236/***.2019.***** 55 Journal of Modern Physics energy scale ∗Λ Remark VI.2.1. Note that in subsection we pointed out that under Assumption VI.1 if bare cosmological constant 0λ = the total cosmological constant vacλ is on at least the order 5 eff nμ − + of magnitude of the renormalized vacuum energy density generated by zero-point fluctuations of standard particle fields and ghost particle fields ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) eff eff 4 2 vac eff 0 4 2 vac eff 0 1 ln d , 8 1 ln d . 8 f O p f O μ μ ε μ μ μ μ μ μ μ μ μ μ − ∗ − ∗ = + Λ = − + Λ ∫ ∫ (287) 6.3. Effective Field Theory and Lorentz Invariance Violation To prevent the vacuum energy density from diverging,the traditional approach also assumes that performing a high-energy cutoff is acceptable. This type of regularization is a common step in renormalization procedures, which aim to eventually arrive at a physical, cutoff-independent result. However, in the case of the vacuum energy density, the result is inherently cutoff dependent, scaling quartically with the cutoff ∗Λ . Remark VI.3.1. By restricting to modes with particle energy a certain cutoff energy ω ∗≤ Λk a finite, regularized result for the energy density can be obtained. The result is proportional to 4∗Λ . Any other fields will contribute similarly, so that if there are bn bosonic fields and fn fermionic fields, the density scales with ( )4b fn n− 4∗Λ . Typically, the cutoff is taken to be near = 1 in Planck units (i.e.the Planck energy), so the vacuum energy gives a contribution to the cosmological constant on the order of at least unity according to Equation (6.2.4). Thus we see the extreme ne-tuning problem: the original cosmological constant λ must cancel this large vacuum energy density vac 1ε  to a precision of 1 in 10120 -but not completelyto result in the observed value 120 eff 10λ −= [5]. Remark VI.3.2. As it pointed out in this paper that a high-energy theory, i.e. QFT in fractal space-time with Hausdorff-Colombeau negative dimension would not display the zero-point fluctuations that are characteristic of QFT, and hence that the divergence caused by oscillations above the corresponding cutoff frequency is unphysical. In this case, the cutoff ∗Λ is no longer an intermediate mathematical construct, but instead a physical scale at which the smooth, continuous behavior of QFT breaks down. Poincaré group of the momentum space is deformed at some fundamental high-energy cutoff ∗Λ The canonical quadratic invariant 2 ab a bp p pη= collapses at high-energy cutoff ∗Λ and being replaced by the non-quadratic invariant: ( ) 2 0 . 1 ab a bp pp l p η ∗Λ = + (288) Remark VI.3.3. In contrary with canonical approach the total effective cosJ. Foukzon et al. DOI: 10.4236/***.2019.***** 56 Journal of Modern Physics mological constant effλ depends only on mass distribution ( )f μ and constant eff effm cμ = but cannot depend on large energy scale ∗Λ . 6.4. Semiclassical Moller-Rosenfeld Gravity Assumption 5 means that it is valid to replace the right-hand side of the Einstein equation Tμν with its expectation Tμν . It requires that either gravity is not in fact quantum, and the Moller-Rosenfeld approach is a complete description of reality, or at least a valid approximation in the weak field limit. The usual argument states that the vacuum state 0 should be locally Lorentz invariant so that observers agree on the vacuum state. This means that the expectation value of the energy-momentum tensor on the vacuum, 0 0Tμν , must be a scalar multiple of the metric tensor gμν which is the only Lorentz invariant rank ( )0,2 tensor. By using Moller-Rosenfeld approach the Einstein field equations of general relativity, a term representing the curvature of spacetime Rμν is related to a term describing the energy-momentum of matter 0 0Tμν , as well as the cosmological constant λ and metric tensor gμν reads: 1 8π 0 0 . 2 R R g g Tυμν υ μν μν μνλ− + = (299) The 00T component is an energy density, we label vac0 0Tμν ε= , so that the vacuum contribution to the right-hand side of Equation (6.4.1) can be written as vac 8π 0 0 8π .T gμν μνε= (290) Subtracting this from the right-hand side of Equation (6.4.1) and grouping it with the cosmological constant term replaces with an "effective" cosmological constant [5]: eff vac8π .λ λ ε= + (291) Note that in flat spacetime, where ( )1, 1, 1, 1g diagμν = − + + + , Eq.(6.4.2) implies vac vacpε = − , where vac 0 0iip T= for any 1, 2,3i = is the pressure. Obviously this implies that if the energy density is positive as is usually assumed, then the pressure must be negative, a conclusion which extends to any metric gμν with a ( )1, 1, 1, 1− + + + signature. Remark VI.4.1. In this paper we assume that the vacuum state 0 should be locally invariant under modified Lorentz boost (1.1.18) but not locally Lorentz invariant. Obviously this assumption violate the Equation (6.4.2). However modified Lorentz boosts (1.1.18) becomes Lorentz boosts for sufficiently small energies and therefore in IR region one obtain in a good approximation vac 8π 0 0 8πT gμν μνε≈ (292) and eff vac8π .λ λ ε≈ + (293) Thus Moller-Rosenfeld approach holds in a good approximation. J. Foukzon et al. DOI: 10.4236/***.2019.***** 57 Journal of Modern Physics 6.5. Quantum Gravity At Energy Scale ∗Λ ≤ Λ . Controllable Violation of the Unitarity Condition Gravitational actions which include terms quadratic in the curvature tensor are renormalizable. The necessary Slavnov identities are derived from Becchi-Rouet-Stora (BRS) transformations of the gravitational and Faddeev-Popov ghost fields. In general, non-gauge-invariant divergences do arise, but they may be absorbed by nonlinear renormalizations of the gravitational and ghost fields and of the BRS transformations [14]. The geneic expression of the action reads ( )4 2 2d 2 ,symI x g R R R Rμνμνα β κ −= − − − +∫ (294) where the curvature tensor and the Ricci is defined by Rλ λμαν ν μα= ∂ Γ and R Rλμν μλν= correspondingly, 2 32πGκ = . The convenient definition of the gravitational field variable in terms of the contravariant metric density reads .h g gμν μν μνκ η= − − (295) Analysis of the linearized radiation shows that there are eight dynamical degrees of freedom in the field. Two of these excitations correspond to the familiar massless spin-2 graviton. Five more correspond to a massive spin-2 particle with mass 2m . The eighth corresponds to a massive scalar particle with mass 0m . Although the linearized field energy of the massless spin-2 and massive scalar excitations is positive definite, the linearized energy of the massive spin-2 excitations is negative definite. This feature is characteristic of higher-derivative models, and poses the major obstacle to their physical interpretation. In the quantum theory, there is an alternative problem which may be substituted for the negative energy. It is possible to recast the theory so that the massive spin-2 eigenstates of the free-fieid Hamiltonian have positive-definite energy, but also negative norm in the state vector space. These negative-norm states cannot be excluded from the physical sector of the vector space without destroying the unitarity of the S matrix. The requirement that the graviton propagator behaves like 4p− for large momenta makes it necessary to choose the indefinite-metric vector space over the negative-energy states. The presence of massive quantum states of negative norm which cancel some of the divergences due to the massless states is analogous to the Pauli-Villars regularization of other field theories. For quantum gravity, however, the resulting improvement in the ultraviolet behavior of the theory is sufficient only to make it renormalizable, but not finite. Remark 6.5.1. (I) The renormalizable models which we have considered in this paper many years mistakenly regarded only as constructs for a study of the ultraviolet problem of quantum gravity. The difficulties with unitarity appear to preclude their direct acceptability as canonical physical theories in locally Minkowski space-time. In canonical case they do have only some promise as phenomenological models. (II) However, for their unphysical behavior may be restricted to arbitrarily large energy scales ∗Λ mentioned above by an appropriate limitation on the J. Foukzon et al. DOI: 10.4236/***.2019.***** 58 Journal of Modern Physics renormalized masses 2m and 0m . Actually, it is only the massive spin-two excitations of the field which give the trouble with unitarity and thus require a very large mass. The limit on the mass 0m is determined only by the observational constraints on the static field. 6.6. Pauli-Villars Masses Distribution Corresponding to Ghost Matter Sector above High Energy Cutoff ∗Λ A Common Origin of the Dark Energy and Dard Matter Phenomena Dark matter is a hypothetical form of matter that is thought to account for approximately 85% of the matter in the universe, and about a quarter of its total energy density. The majority of dark matter is thought to be non-baryonic in nature, possibly being composed of some as-yet undiscovered subatomic particles. Its presence is implied in a variety of astrophysical observations, including gravitational effects that cannot be explained unless more matter is present than can be seen. For this reason, most experts think dark matter to be ubiquitous in the universe and to have had a strong influence on its structure and evolution. Dark matter is called dark because it does not appear to interact with observable electromagnetic radiation, such as light, and is thus invisible to the entire electromagnetic spectrum, making it extremely difficult to detect using usual astronomical equipment [36] [37] [38]. Figure 8 Analysis of a giant new galaxy survey, made with ESO's VLT Survey Telescope in Chile, suggests that dark matter may be less dense and more smoothly distributed throughout space than previously thought. An international team used data from the Kilo Degree Survey (KiDS) to study how the light from about 15 million distant galaxies was affected by the gravitational influence of matter on the largest scales in the Universe. The results appear to be in disagreement with earlier results from the Planck satellite. This map of dark matter in the Universe was obtained from data from the KiDS survey, using the VLT Survey Telescope at ESO's Paranal Observatory in Chile. It reveals an expansive web of dense (light) and empty (dark) regions. This image is one out Figure 8. Dark matter map for a patch of sky based on gravitational lensing analysis [38] Hilderbrandt 16. J. Foukzon et al. DOI: 10.4236/***.2019.***** 59 Journal of Modern Physics of five patches of the sky observed by KiDS. Here the invisible dark matter is seen rendered in pink, covering an area of sky around 420 times the size of the full moon. This image reconstruction was made by analyzing the light collected from over three million distant galaxies more than 6 billion light-years away. The observed galaxy images were warped by the gravitational pull of dark matter as the light traveled through the Universe. Some small dark regions, with sharp boundaries, appear in this image. They are the locations of bright stars and other nearby objects that get in the way of the observations of more distant galaxies and are hence masked out in these maps as no weak-lensing signal can be measured in these areas [38]. The luminous (light-emitting) components of the universe only comprise about 0.4% of the total energy. The remaining components are dark. Of those, roughly 3.6% are identified: cold gas and dust, neutrinos, and black holes. About 23% is dark matter, and the overwhelming majority is some type of gravitationally self-repulsive dark energy. Remark 6.6.1. In order to explain physical nature of the dark matter sector we assume that the main part of dark matter, i.e., 23% 4.6% 18%− = (see Figure 9) formed by supermassive ghost particles with masses such that 2mc ∗> Λ , see ref. [12], Subsection 2.3. Remind that vacuum energy density for free scalar quantum field with a wrong statistic is: ( ) ( ) ( )2 2 2 2 2 23 0 0 1 1 4π d d , 2 2π c p p p K p p p K Iε μ μ μ μ ∞ ∞ ′ ′= − + = + =∫ ∫  (296) where mcμ = . From the basic definitions [1] ( ) ( ) 2 3 0 2 2 1 1 1, 4π d , , , 2 32π xx x x x x cp T p u p p p u p p μ μ ∞ = = − = = + ∫  pu u p (297) one obtains ( ) ( ) 4 0 2 2 d . 3 K p pp K F p μ μ μ ∞′ ′= = + ∫ (298) Remark 6.6.2. Note that the integral in RHS of Equation (297) and in Equation (298) is divergent and ultraviolet cutoff is needed. Thus in accordance with [1] we set ( ) ( ) ( ) ( )0 0 0 0, , , , , ,p K I p p p K F pε μ μ μ μ′ ′= = (299) where ( ) ( )0 0 4 2 2 2 0 00 0 2 2 d, d , , , p p p pI p p p p F p p μ μ μ μ = + = + ∫ ∫ (300) where 0p c∗≤ Λ . For fermionic quantum field with a wrong statistic, similarly one obtains ( ) ( ) ( ) ( )0 0 0, 4 , , 4 , .p K I p p K F pε μ μ μ μ′ ′= − = − (301) Thus from Equations. (300)-(301) by using formally Pauli-Villars regularization [7] [8] and regularization by high-energy cutoff the expression for J. Foukzon et al. DOI: 10.4236/***.2019.***** 60 Journal of Modern Physics Figure 9. Matter and energy distribution in the universe today. The luminous (light-emitting) components of the universe only comprise about 0.4% of the total energy. The remaining components are dark. free vacuum energy density ε reads ( ) 2 vac 0 0 , M i i i f I pε μ = = ∑ (302) and the expression for pressure p reads ( ) 2 vac 0 0 , . M i i i p f F pμ = = ∑ (303) Definition 6.6.1. We define now discrete distribution :PVf + →  by formula ( ) ,PV i if fμ = (304) and we will call it as a full discrete Pauli-Villars masses distribution. Remark 6.6.3. We assume now that in Equations (302)-(303): 1) the quantities . , 1, 2, ,s mi i i Mμ μ= =  are masses of physical particles corresponding to standard matter and 2) the quantities . , 1, 2, , 2g mi i i M Mμ μ= = +  are masses of ghost particles with a wrong kinetic term and wrong statistics corresponding to physical dark matter. Remark 6.6.4. We recall that the Euler-Maclaurin summation formula reads ( )( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 2 1 2 11 1 2 1 2 2 2 1 1 d 1, . M M i M M g i h f A g g A h g g O h f g h μ μ μ μ μ μ μ μ μ μ μ = + − = + −   ′ ′+ − + =   ∑ ∫ (305) Let ( )g μ be an appropriate continuous function such that: 1) ( ) , 1, 2, , 2i ig f i Mμ = =  , 2) ( ) ( )2 10, 0Mg gμ μ′ ′= = . Thus from Equations (302)-(303) and Equations (305) we obtain ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 2 1 2 vac 0 0 0 2 1 2 2 0 1 1 0 , , d , , M M i i i M M f I p f I p A h f I p f I p O h μ μ ε μ μ μ μ μ μ μ μ = = =  + − +  ∑ ∫ (306) J. Foukzon et al. DOI: 10.4236/***.2019.***** 61 Journal of Modern Physics and ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 2 1 2 vac 0 0 0 2 1 2 2 0 1 1 0 , , d , , . M M i i i M M p f F p f F p A h f F p f F p O h μ μ μ μ μ μ μ μ μ μ = = =  + − +  ∑ ∫ (307) Definition 6.6.2. We will call the function ( )PVf μ as a full continuous Pauli-Villars masses distribution. Definition 6.6.3. We define now: 1) discrete distribution . . :b g mPVf + →  by formula ( ). . . , 1, 2, ,b g m s mPV i if f i Mμ = =  (308) and we will call it as discrete Pauli-Villars masses distribution of the bosonic ghost matter and 2) discrete distribution . . :f g mPVf + →  by formula ( ). . , 1, 2, , 2f g mPV i if f i M Mμ = = +  (309) and we will call it as discrete Pauli-Villars masses distribution of the fermionic ghost matter. Remark 6.6.4. We rewrite now Equations (306)-(307) in the following equivalent form ( ) ( ) ( )( ) ( )( ) ( ) 2 . . . . . . . . . . . vac 0 0 1 = 1 , , M M b g m s m b g m f g m f g m f g m PV i i PV j i j i i j i M f I p f I pε μ μ μ μ = + = +∑ ∑ (310) and ( ) ( ) ( )( ) ( )( ) ( ) 2 . . . . . . . . . . . . vac 0 0 1 1 , , , M M b g m b g m b g m f g m f g m f g m PV i i PV j i j i i j i M p f F p f F pμ μ μ μ = = + = +∑ ∑ (311) where ( ) , 1 1,2, ,j i i M i M= + = +  . Remark 6.6.6. We assume now that:1) ( ) . . . .b g m f g m i j iμ μ≈ , 2) ( ) ( )( ). . . . . . . . 1b g m b g m f g m f g mPV i PV j if fμ μ+  , i.e., ( ) ( )( ). . . . . . . . .b g m b g m f g m f g mPV i PV j if fμ μ≈ − (312) Note that Equation (312) meant highly symmetric discrete Pauli-Villars masses distribution between bosonic ghost matter and fermionic ghost matter above that scale .∗Λ Thus from Equations (310)-(311) and Equations (312) we obtain ( ) ( ) ( )( ) ( )( ) ( ) ( ) ( )( ) ( ) 2 . . . . . . . . . . . . vac 0 0 1 1 . . . . . . . . 0 1 , , , M M b g m b g m b g m f g m f g m f g m PV i i PV j i j i i j i M M b g m b g m f g m f g m PV i PV ij i i f I p f I p f f I p ε μ μ μ μ μ μ μ = = + = = +  = +  ∑ ∑ ∑ (313) and ( ) ( ) ( )( ) ( )( ) ( ) ( ) ( )( ) ( ) 2 . . . . . . . . . . . . vac 0 0 1 1 . . . . . . . . 0 1 , , , . M M b g m b g m b g m f g m f g m f g m PV i i PV j i j i i j i M M b g m b g m f g m f g m PV i PV ij i i p f F p f F p f f F p μ μ μ μ μ μ μ = = + = = +  = +  ∑ ∑ ∑ (314) J. Foukzon et al. DOI: 10.4236/***.2019.***** 62 Journal of Modern Physics From Equations (313)-(314) and Equations (305) finally we obtain ( ) ( )( ) ( ) ( ) ( ) ( )eff 1 . . . . . . . . vac 0 1 . . . . 0 , , d M b g m b g m f g m f g m PV i PV ij i i b g m f g m PV PV f f I p f f I p μ μ ε μ μ μ μ μ μ μ =  = +   = +  ∑ ∫ (315) and ( ) ( )( ) ( ) ( ) ( ) ( )eff . . . . . . . vac 0 1 . . . . 0 1 , , d , M b g m s m f g m f g m PV i PV ij i i b g m f g m PV PV p f f F p f f F p μ μ μ μ μ μ μ μ μ =  = +   = +  ∑ ∫ (316) where obviously ( ) ( ) ( ). . . . . . 0.b g m f g m g mPV PV PVf f fμ μ μ+ = ≈ (317) Thus finally we obtain ( ) ( )( ) ( ) ( )( ) ( )2 eff 1 eff 1 2. . . . eff eff 0 0, , , d , g m g m PVp f I p μ μ ε μ μ μ μ μ= ∫ (318) and ( ) ( )( ) ( ) ( )( ) ( )2 eff 1 eff 1 2. . . . eff eff 0 0, , , d , g m g m PVp p f F p μ μ μ μ μ μ μ= ∫ (319) where ( ) ( )1 2eff eff 0, pμ μ  . In order to calculate ( ) ( )( )1 2. . eff eff 0, ,g m pε μ μ and ( ) ( )( )1 2. . eff eff 0, ,g mp pμ μ let us evaluate now the following quantities defined above by Equations (300) ( ) 0 0 2 2 2 2 2 0 2 0 0 , d 1 d p p pI p p p p p pμ μ μ μ = + = +∫ ∫ (320) and ( ) 0 04 4 1 0 2 2 2 0 0 2 1 d 1 d, , 3 3 1 p pp p p pF p p p μμ μ μ − = = + + ∫ ∫ (321) where 0 1p μ  . Note that 2 2 4 6 2 2 4 6 2 2 4 6 2 2 2 2 2 2 2 4 6 4 6 8 2 3 5 1 1 11 1 2 8 16 1 1 11 1 2 8 16 1 1 1 2 8 16 p p p p p p p pp p p p p p pp μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ + = + − + +   + = + = + − + +    = + − + +    (322) By inserting Equation (322) into Equation (320) one obtains ( ) 0 4 6 8 2 0 3 5 0 5 7 9 3 0 0 0 0 3 5 1 1 1, d 2 8 16 1 1 1 1 3 10 7 8 9 16 p p p pI p p p p p p p μ μ μ μ μ μ μ μ μ   = + − + +    = + − + + × × ∫   (323) Note that J. Foukzon et al. DOI: 10.4236/***.2019.***** 63 Journal of Modern Physics 1 22 2 4 2 2 4 1 22 4 6 8 4 1 2 3 5 1 31 1 2 8 1 31 2 8 p p p p p p pp μ μ μ μ μμ μ μ − − −   + = − + +      + = − + +      (324) By inserting Equation (324) into Equation (321) one obtains ( ) 0 04 1 4 6 8 0 3 52 0 0 2 5 7 9 0 0 0 3 5 1 d 1 1 3, d 3 3 2 8 1 1 1 3 5 2 3 7 8 9 p pp p p p pF p p p p p p μμ μ μ μ μ μ μ μ −   = = − + +   + = − + + × × × × ∫ ∫   (325) By inserting Equation (323) into Equation (318) one obtains ( ) ( )( ) ( ) ( )( ) ( ) ( )( ) ( ) ( )( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 2 eff 1 eff 2 eff 1 eff 2 2 2 eff eff eff 1 1 1 eff eff eff 1 2. . eff eff 0 . . 0 5 7 9 . . 3 0 0 0 0 3 5 . . . .3 5 7 . .0 0 0 3 , , , d 1 1 1 1 d 3 10 7 8 9 16 d d d 3 10 7 8 g m g m PV g m PV g m g m PV PVg m PV p f I p p p p f p f fp p p f μ μ μ μ μ μ μ μ μ μ ε μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ =   = + − + +  × ×  = + − + × ∫ ∫ ∫ ∫ ∫   (326) By inserting Equation (325) into Equation (319) one obtains ( ) ( )( ) ( ) ( )( ) ( ) ( )( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) ( ) 2 eff 1 eff 2 eff 1 eff 2 2 2 eff eff eff 1 1 1 eff eff eff 1 2. . eff eff 0 . . 0 5 7 9 . . 0 0 0 3 5 . . . . . .5 7 9 0 0 0 3 5 , , , d 1 1 d 3 5 2 3 7 8 9 d d d 3 5 2 3 7 8 9 g m g m PV g m PV g m g m g m PV PV PV p p f F p p p p f f f fp p p μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ μ =   = − + +  × × × ×  = − + + × × × × ∫ ∫ ∫ ∫ ∫   (327) Remark 6.6.7. We assume now that ( ) ( )( ) ( ) ( ) ( ) 1 1 2 eff eff eff. . 2 eff , 7 0 n g m PV O n f μ μ μ μ μ μ μ −   > ≤ ≤   =   > (328) Note that under assumption (328) the quantities ( ) ( )( )1 2. . eff eff 0, ,g m pε μ μ and ( ) ( )( )1 2. . eff eff 0, ,g mp pμ μ cannot contribute in the value of the cosmological constant. 7. Conclusion We argue that a solution to the cosmological constant problem is to assume that there exists hidden physical mechanism which cancels divergences in canonical 4 4,QED QCD , Higher-Derivative-Quantum-Gravity, etc. In fact, we argue that corresponding supermassive Pauli-Villars ghost fields, etc. really exist. New J. Foukzon et al. DOI: 10.4236/***.2019.***** 64 Journal of Modern Physics theory of elementary particles which contain hidden ghost sector is proposed. In accordance with Zel'dovich hypothesis [1] we suggest that physics of elementary particles is separated into low/high energy ones, the standard notion of smooth spacetime is assumed to be altered at a high energy cutoff scale ∗Λ and a new treatment based on QFT in a fractal spacetime with negative dimension is used above that scale. This would fit in the observed value of the dark energy needed to explain the accelerated expansion of the universe if we choose highly symmetric masses distribution below that scale ∗Λ , i.e., ( ) ( ) 2. . eff eff, ,s m g mf f cμ μ μ μ μ ∗≤ < Λ Acknowledgements We thank the reviewers for their comments. Conflicts of Interest The authors declare no conflicts of interest regarding the publication of this paper. References [1] Zel'dovich, Ya.B. (1968) The Cosmological Constant and the Theory of Elementary Particles. Soviet Physics Uspekhi, 11, 381-393. https://doi.org/10.1070/PU1968v011n03ABEH003927 [2] Gliner, É.B. 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