De Witt convention [66]: ω ¼ðd − 4Þðd þ 1Þ=8. In this
case, the local factor does not contribute when the fixed-
dimension renormalization in d ¼ 4 is employed.
The sum in (9) is a sum over four topologies, that is, the
sum over topologically distinct manifolds Σ
i
(analogue to
the sum over genera in string theory or the sum over
homotopically inequivalent vacua in the Yang-Mills
theory), which can potentially contain topological phase
factors, e.g., the Euler-Poincar´e characteristic of Σ
i
,
cf. Refs. [67].
2
For future convenience, we note that the R
2
part in the
Weyl action S in (6) can be further decomposed with the
help of the Hubbard-Stratonovich (HS) transformation
[10,69–71] as
expðiS
R
Þ ≡ exp
i
6α
2
Z
d
4
x
ffiffiffiffiffi
jgj
p
R
2
¼
Z
Dϕ exp
−i
Z
d
4
x
ffiffiffiffiffi
jgj
p
ϕR þ
3
8ω
C
ϕ
2
:
ð10Þ
It is not difficult to see that the essence of the HS
transformation (10) is a straightforward manipulation of
a functional Gaussian integral (shifting the quadratic
trinomial in the exponent). Although an auxiliary HS field
ϕðxÞ does not have a bare kinetic term, one might expect
that due to radiative corrections it will develop in the IR
regime a gradient term which will then allow us to identify
the HS boson with a genuine propagating mode. This
scenario is, in fact, well-known from the condensed matter
theory. A quintessential example of this is obtained when
the BCS superconductivity is reduced to its low-energy
effective level. There, the HS boson coincides with the
disordered field whose dynamics is described via the
celebrated Ginzburg-Landau equation [71,72].
The ϕ field can be separated into a background field hϕi
corresponding to a VEV of ϕ plus fluctuations δϕ. Since
hϕi is dimensionful, it must be zero in the case when the
theory is scale invariant. On the other hand, when the scale
invariance is broken, ϕ will develop a nonzero VEV. So, the
HS field ϕ plays the role of the order-parameter field. The
inner workings of this mechanism were illustrated in
Ref. [10], where it was shown than on the flat background
ϕ develops (in the broken phase) a nonzero VEV. With the
benefit of hindsight, we further introduce an arbitrary
“mixing” hyperbolic angle ϑ ∈ ð−∞; ∞Þand write formally
S
R
¼ S
R
cosh
2
ϑ − S
R
sinh
2
ϑ: ð11Þ
Applying now the HS transformation to the S
R
sinh
2
ϑ part,
we get
S
R
¼ −
Z
d
4
x
ffiffiffiffiffi
jgj
p
ϕR þ
2ω
C
cosh
2
ϑ
3
Z
d
4
x
ffiffiffiffiffi
jgj
p
R
2
þ
3
8ω
C
sinh
2
ϑ
Z
d
4
x
ffiffiffiffiffi
jgj
p
ϕ
2
: ð12Þ
Although the full theory described by the action S is
independent of the mixing angle ϑ, truncation of the
perturbation series after a finite loop order in t he fluctu-
ating metric field will destroy this independence. The
optimal result is reached by employing the principle of
minimal sensitivity [73,74] known from the RG calculus.
There, if a perturbation theory depends on some unphys -
ical parameter (as ϑ in our case), the best result is achieved
if each order has the weakest possible dependence on the
parameter ϑ. Consequently, at each loop order, the value o f
ϑ is determined from the vanishing of th e corresponding
derivative of effective acti on.
As discussed in [10], the fluctuations of the metric g
μν
can make hϕinot only nonzero but one can also find a set of
parameters in a model’s parameter space for which
hϕi ∼ M
2
P
where M
P
¼ 2.44 × 10
18
GeV is related to the
Planck energy scale. Consequently, Newton’s constant κ
N
is dynamically generated. Owing to the last term in (12), the
existence of dynamical dark energy (a dynamical cosmo-
logical constant) is also an automatic consequence of the
theory. In addition, by assuming that in the broken phase a
cosmologically relevant metric is that of the Friedmann-
Lemaître-Robertson-Walker (FLRW) type, then, modulo a
topological term, the additional constraint,
Z
d
4
x
ffiffiffiffiffi
jgj
p
3R
2
μν
¼
Z
d
4
x
ffiffiffiffiffi
jgj
p
R
2
; ð13Þ
holds due to a conformal flatness of the FLRW metric
[75,76]. It was argued in [10] that from (6) and (12) one
obtains in the broken phase the effective gravitational
action of the form,
S ¼ −
1
2κ
2
N
Z
d
4
x
ffiffiffiffiffi
jgj
p
ðR − ξ
2
R
2
− 2Λ
cc
Þ; ð14Þ
where both κ
N
(Newton’s constant) and ξ (Starobinsky’s
parameter) are dynamically generated. Note that ξ has the
dimension of an inverse mass and by the Planck satellite
data ξ=κ
N
∼ 10
5
(cf. Ref. [1]). The action (14) is nothing
but the Starobinsky action with the cosmological constant.
We stress that the cosmological constant Λ
cc
is entirely of
2
The sum over four topologies is a problematic concept since
four manifolds are generally unclassifiable—that is, there is no
algorithm that can determine whether two arbitrary four mani-
folds are homeomorphic. On the other hand, simply connected
compact topological four manifolds are classifiable in terms of
Casson handles shown by M. H. Freedman [68], which can be
applied in functional integrals in Euclidean gravity. In the
Lorentzian case, one simply restricts oneself to some subset of
four manifolds. If this subset is closed under a composition of the
functional integral, then a theory thereby obtained is at least
naively self-consistent.
INFRARED BEHAVIOR OF WEYL GRAVITY: FUNCTIONAL … PHYS. REV. D 101, 044050 (2020)
044050-5