The Renormalization Group flow of unimodular f(R) gravity Astrid Eichhorn Blackett Laboratory, Imperial College, Prince Consort Road, London SW7 2AZ, United Kingdom E-mail: [email protected] Unimodular gravity is classically equivalent to General Relativity. This equivalence extends to actionswhicharefunctionsofthecurvaturescalar. Atthequantumlevel,thedynamicscoulddiffer. Most importantly, the cosmological constant is not a coupling in the unimodular action, providing a new vantage point from which to address the cosmological constant fine-tuning problem. Here, a quantumtheorybasedontheasymptoticsafetyscenarioisstudied,andevidenceforaninteracting fixed point in unimodular f(R) gravity is found. We study the fixed point and its properties, and also discuss the compatibility of unimodular asymptotic safety with dynamical matter, finding evidence for its compatibility with the matter degrees of freedom of the Standard Model. 5 I. INTRODUCTION 1 0 2 There are only very few observations from which one could expect to learn something about the deep structure of spacetime, described by a model of quantum gravity. One of those is the observed accelerated expansion of the r p universe, which can be modelled by including a nonzero cosmological constant in the Einstein-Hilbert action. This A entailsthecosmologicalconstantproblem. Here,wewillfocusononeaspectofthisproblem,namelythequestionwhy quantumvacuumfluctuationsdonotseemtogravity, i.e., whythecosmologicalconstantexhibitsaseverefine-tuning 1 problem. Asitsmass-dimensionalityis2,onewouldexpectquantumfluctuationstodriveittobeoforderoneinunits defined by the square of the physical mass scale of the theory, which is the Planck scale. In terms of Renormalization ] c Group (RG) trajectories, the tiny value inferred from observations implies that a particularly fine-tuned trajectory q has to be picked. Of course every so-called relevant coupling in a quantum theory corresponds to one free parameter - r that can only be fixed by comparison with experiment. Thus one always has to pick a particular RG trajectory in g order for the model to reproduce observations. On the other hand, relevant couplings with only, e.g., a logarithmic [ running imply that if one picks a trajectory nearby, the measured value of the coupling will only change slightly. For 2 the cosmological constant, this statement is not true, i.e., for a reason which as yet has no dynamical explanation, v our universe just ”happens” to live on a highly fine tuned choice of trajectory. 8 As already proposed by Weinberg [1], a ”degravitation” of the cosmological constant is possible by changing its 4 status from a coupling in the action to a constant of integration that arises at the level of the equations of motion. 8 The second is a ”classical” quantity in the sense that it is not affected by quantum fluctuations. In unimodular 5 √ gravity [2], the metric is conceived as a symmetric tensor with fixed determinant −g = (cid:15) [3]. This implies that no 0 √ . operatoroftheform −g exists,asthevolumeisjustafixednumber,andthecosmologicalconstantisremovedfrom 1 the space of couplings, the theory space. Once a Renormalization Group trajectory in this reduced space has been 0 picked, the effective equations of motion can be calculated from the full effective action – the infrared endpoint of the 5 1 trajectory – and the cosmological constant will then make its appearence as a constant of integration. In this way, : quantum vacuum fluctuations do not affect the value of the cosmological constant. It is thus of interest to investigate v a quantum theory of unimodular gravity. Unimodular gravity in both its quantum and classical form has sparked i X considerable interest since it was originally proposed [3–9]. r A second motivation to consider unimodular quantum gravity lies in the fact it will most probably differ from a the non-unimodular version of quantum gravity. This inequivalence arises, as imposing the unimodularity condition alters the spectrum of fluctuations of the theory. In more detail, deriving the full metric propagator by taking the √ second variation of the action yields different results when −g = (cid:15) is imposed, than if the metric determinant is allowed to fluctuate. In particular, fluctuations of the conformal mode, which yield an instability of the path-integral in the Euclidean case when starting from the Einstein-Hilbert action, are absent in unimodular gravity. This already suggests that although classically equivalent [6], the quantum version of unimodular and ”standard” gravity could differ. The absence of the conformal instability even suggests that the unimodular quantum theory could have better properties. Asthesearchforthe”quantumtheoryofgravitydescribingouruniverse”isstillongoing,anexplorationofdifferent models for quantum gravity models is clearly of interest, both from a theoretical as well as from a phenomenological point of view. Here we will focus on exploring models of asymptotically safe quantum gravity. As we will point out in Sec. V, it might be possible to distinguish between different versions of asymptotically safe gravity experimentally. Finally, inordertobetterunderstandthestructureofRenormalizationGroup(RG)flowsingravityitisofinterest to consider settings with fewer propagating degrees of freedom in the path-integral. Here, we should clarify that the physically propagating degrees of freedom in both settings, unimodular vs. ”full” gravity, agree in a perturbative 2 expansionaroundaflatbackground,i.e.,thereisamasslessspin-2graviton[6]. Ontheotherhand,theconfigurations that enter the path integral are different metric configurations, and, e.g., the configuration space of the conformal modeisalsosummedoverinthecaseof”full”gravity. Inordertoshedlightonthephysicalmechanismofasymptotic safety, it is helpful to consider settings where some of the modes in the path-integral are removed. In this paper, we will consider unimodular asymptotic safety, and investigate truncated Renormalization Group flows based on an f(R) action. We will discuss the classical equivalence of ”full” gravity, which we will call Einstein gravity, with unimodular gravity based on an f(R) action in Sec. IIA. We will then focus on the quantum theory, and review the asymptotic safety scenario and the functional RG in Sec. III, where we will also present all technical details of our calculation. In Sec. IV we will present the flow equation for f(R) and discuss a fixed point and its properties. We will make a first step toward phenomenology in studying the effect of dynamical matter in Sec. V, and finally conclude in Sec. VI. II. RELATION BETWEEN UNIMODULAR GRAVITY AND EINSTEIN GRAVITY A. Classical equivalence of f(R) gravity and unimodular f(R) Before we embark on an analysis of quantum gravity, let us clarify the classical relation between unimodular f(R) gravity and f(R) gravity with a full metric, see, e.g., [10] for a review. Here we will focus on the Lorentzian case, and then switch to a Euclidean setting for the analysis of the quantum theory. We focus on actions given by a function of the curvature scalar, f(R), with f(0) = 0. We introduce the Newton coupling G and the cosmological constant Λ N in the action (cid:90) √ (cid:18) 1 (cid:19) S = d4x −g f(R)+ Λ+L . (1) 8πG m N The corresponding equations of motion are given by 1 1 1 − f(R)g +f(cid:48)(R)R −D D f(cid:48)(R)+g D2f(cid:48)(R)+ Λg = T , (2) 2 µν µν µ ν µν 16πG µν 2 µν N where the energy-momentum tensor is given by 2 δL T =−√ m. (3) µν −gδgµν The Bianchi-identities (cid:18) (cid:19) 1 Dµ R − g R =0, (4) µν 2 µν will now play a crucial role: By taking the covariant derivative of Eq. (2), we deduce the conservation law for the energy-momentum tensor by imposing the Bianchi-identities. To obtain the unimodular equations of motion for the action (cid:90) S = d4x(cid:15)(f(R)+L ), (5) u m we have to consider tracefree variations gµνδg =0 and obtain µν (cid:18) (cid:19) 1 1 1 1 f(cid:48)(R)R −D D f(cid:48)(R)+ g D2f(cid:48)(R)− g Rf(cid:48)(R)= T − g Tλ . (6) µν µ ν 4 µν 4 µν 2 µν 4 µν λ Crucially, the covariant derivative of the lhs of Eq. (6) does not vanish when the Bianchi-identities are used. Instead, we can impose conservation of the energy-momentum tensor1, and thereby derive a nontrivial identity, namely (cid:18) (cid:19) 1 3 1 −1 D − f(cid:48)(R)R− D2f(cid:48)(R)+ Tλ = D f(R), (7) ν 4 4 8 λ 2 ν 1 As discussed, e.g., in [5], it is not clear whether this requirement can be preserved in a quantum field theory setting for the matter degreesoffreedom. Thiscouldpotentiallyleadtoasituationwherethelow-energyeffectiveequationsofmotionallowustodistinguish betweenGeneralRelativityandunimodulargravity. 3 where we have used Eq. (6) and imposed Eq. (4). This identity allows us to identify 1/2f(R) with −1f(cid:48)(R)R− 4 3D2f(cid:48)(R)+1Tλ,uptoaconstantofintegration,whichwechoosetocall 1 Λ. InsertingthisidentityintoEq.(6) 4 2 λ 16πGN weobtainEq.(2), i.e., classicallyunimodularf(R)gravitycannotbedistinguishedfromstandardf(R)gravity. Note that this statement depends on the postulate of energy-momentum conservation in the unimodular case. It is a priori clearthatthetwotheoriescanonlybeclassicallyequivalent, ifoneadditionalconditionisimposedintheunimodular case: Since the equations of motion of unimodular gravity are obtained by removing the trace from the standard equations of motion, they contain precisely one condition less. III. UNIMODULAR QUANTUM GRAVITY A. Asymptotic safety From now on we will focus on Euclidean quantum gravity, as this allows for a straightforward application of RG tools. Toarriveataunimodularquantumtheoryofgravity,wewillinvoketheasymptoticsafetyconjectureforgravity [11]. Interestingly, unimodularityplaysaroleinanumberofotherapproachestoquantumgravity, e.g.,withincausal set quantum gravity, where a discrete version of unimodularity would be implemented by performing the path-sum over all causets with a fixed number of elements [12, 13]. Further, Causal Dynamical Triangulations is based on a setting where the number of simplices usually is held fixed for the simulations, i.e., the cosmological constant is removed from the space of couplings, see, e.g., [14]. An asymptotically safe quantum theory of gravity is valid at arbitrarily high momenta, i.e., beyond the regime of validityofeffectivefieldtheory[15,16],andatthesametimeremainspredictive,i.e.,onlycomeswithafinitenumber of free parameters. It is the second requirement that breaks down in a perturbative quantization of gravity when extended to arbitrarily high momenta [17–19]. Then, an infinite number of counterterms needs to be introduced and thereisnomechanismtopredictthecorrespondingfreeparameters. Inasymptoticsafety,theRenormalizationGroup flowapproachesaninteractingfixedpointathighmomenta. Specifically,werefertothedimensionlesscouplingshere, whichcanbeobtainedfromthedimensionfulonesbyanappropriaterescalingwiththeRGmomentumscalek. Ifthese couplings approach a fixed point, the theory becomes scale free and can be extended to arbitrarily high momentum scales. The RG flow lives in an infinite dimensional space of all couplings, the theory space, as quantum fluctuations generically generate all operators that are compatible with the symmetries. In this infinite-dimensional space one must then investigate whether predictivity can be obtained, i.e., whether the model has only a finite number of free parameters. This is ensured, if the interacting fixed point comes with a finite number of relevant, i.e., ultraviolet (UV) attractive directions. The (ir)relevance of a coupling g (k) determines its scale-dependence in the vicinity of a i fixed point. This can be obtained after linearising the RG flow: (cid:88) (cid:18) k (cid:19)−θI g (k)=g + C VI . (8) i i∗ I i k 0 I Herein, g denotes the fixed-point values of the coupling g . VI are the eigenvectors of the stability matrix M = i∗ i ij (∂β /∂g )| , and −θ its eigenvalues. C are constants of integration. If θ < 0, then C = 0 is required in gi j gn=gn∗ I I J J order for the couplings to approach the fixed point in the UV limit, where k → ∞. On the other hand, relevant directions are those with θ > 0. These approach the fixed point automatically, imposing no requirement on the J corresponding parameter C . They have instead to be determined experimentally. k is an arbitrary reference scale, J 0 whichcanbetakenasthescaleatwhichthevalueofthecouplingsismeasured. AtaGaußianfixedpoint,suchasthat underlyingasymptoticfreedominYang-Millstheory,onlythecouplingsofpositiveandvanishingmassdimensionality can be relevant. At an interacting fixed point, quantum fluctuations shift the critical exponents away from the mass dimensionality by an anomalous dimension. Additionally, the (ir)relevant directions are typically no longer given by the operators defined at the Gaußian fixed point, but in fact correspond to mixtures of these. When the anomalous dimensions remain finite, this suggests that only a finite number of relevant directions exist, and that these can be found among appropriate combinations of the couplings with the largest mass dimensionalities. It remains for us to determine whether an interacting fixed point exists in unimodular gravity and which of the couplings correspond to relevant ones. Most importantly, we cannot draw any strong conclusions from the evidence forafixedpointinthecaseofquantumEinsteingravity[20–28]. Asthesymmetrychangesfromfulldiffeomorphisms to transverse diffeomorphisms, and the field content is restricted by the unimodularity requirement, the two theory spaces are different. The existence and properties of possible fixed points can therefore be different in the two cases, as discussed in [29]. 4 B. Comments on the equivalence between quantum Einstein gravity and unimodular quantum gravity Inthefollowing,wewillrefertothequantumtheorywithafulldynamicalmetricandfulldiffeomorphismsymmetry as quantum Einstein gravity, also in cases where the underlying action is not the Einstein-Hilbert action. In the √ unimodular case, g =(cid:15) implies that the volume term is not an operator any more, but simply a number. As such, it will be dropped from the action, and no quantum fluctuations contribute to the running of the prefactor, i.e., the cosmological constant. In other words, the cosmological constant is removed from the unimodular theory space. In principle, one could also attempt to impose the unimodularity condition employing a Lagrange multiplier, in which case the cosmological constant would remain a running coupling, presumably resulting in an inequivalent quantum √ model [8]. In quantum Einstein gravity, one can impose g = (cid:15) as a gauge. This does of course not remove the cosmological constant from the theory space. This becomes particularly important when the cosmological constant correspondstoarelevantcouplingatanRGfixedpoint. Thenthepredictedvaluesofallothercouplingsintheinfrared will depend on the corresponding free parameter, associated to the cosmological constant. On the other hand, in the unimodular case, the value of the cosmological constant, which will first enter the theory at the level of the equations of motion, is also a free parameter. Howevernone of the predictable values of the irrelevant couplings depend on that √ parameter. Furthermore, as pointed out in [29], choosing g = (cid:15) as a gauge in quantum Einstein gravity implies that a corresponding Faddeev-Popov ghost contributes to the Renormalization Group flow. This additional ghost is completelyabsentinunimodulargravity. Imposingunimodularitydirectlyontheallowedconfigurationsofthemetric in the path-integral also changes the symmetry from diffeomorphisms to transverse diffeomorphisms; again implying differences at the quantum level. Choosing a classically equivalent formulation of unimodularity can allow to keep full diffeomorphism invariance [7], however we focus on the other case here. √ Finally, imposing g = (cid:15) in unimodular gravity also implies that the spectrum of quantum fluctuations, i.e., the off-shell part of the propagator of metric fluctuations, differs. We will again see this explicitly in Sec. IIIC2. By itself, this already changes the RG flow. We conclude that an equivalence between unimodular quantum gravity and quantum Einstein gravity is not be expected. C. Deriving the flow equations Inthissection,wewilldetailthederivationoftheRGflowequationforunimodularquantumgravity. Wewillfocus on a setting where unimodularity is implemented as a restriction on the allowed configurations in the path integral. This has the advantage that it reduces the number of fluctuating gauge degrees of freedom in the path integral, and could therefore be expected to yield better results already in simple approximations of the full path-integral. 1. Wetterich equation To examine whether asymptotic safety is realized in unimodular gravity, knowledge of the non- perturbative beta functions is required. Here, we will use the framework of the functional Renormalization Group to obtain these. In that setting, an infrared cutoff function R (−D2), (with D denoting a placeholder for the appropriate covariant k Laplacian) depending on the momentum scale k, is included in the generating functional, which suppresses quantum fluctuations with covariant momenta −D2 <k2 [30, 31]. For instance, a theta-cutoff takes the form [33] R (−D2)=(k2−(−D2))θ(k2−(−D2)). (9) k ThisallowsustoobtainthescaledependenteffectiveactionΓ ,whichencodestheeffectofhigh-momentumquantum k fluctuations. Its scale dependence is governed by the Wetterich-equation 1 (cid:16) (cid:17)−1 ∂ Γ = STr Γ(2)+R ∂ R , ∂ =k∂ . (10) t k 2 k k t k t k Herein Γ(2) is the second functional derivative with respect to the quantum fields, which is matrix-valued in field k space. The supertrace implies a summation in field space, including a negative sign for Grassmann-valued fields. It also encodes a summation/integration over the discrete/continuous eigenvalues of the kinetic operator Γ(2). For k reviews, see, e.g., [34]. In the case of gravity, M. Reuter has pioneered the application of the Wetterich equation in [25]. Gravity-specific reviews of the asymptotic safety scenario and the application of the functional Renormalization Group can be found in, e.g., [35]. The application of Renormalization Group methods requires us to set a scale, which seems a challenging task in quantum gravity, where no background spacetime is assumed to exist. Here, the background field method [36] can 5 be employed to circumvent this problem: Splitting the full metric into a background and a fluctuation piece provides a background covariant derivative. This can be used to define ”high-momentum” and ”low-momentum” quantum fluctuations by decomposing the quantum field into eigenfunctions of the background Laplacian and sorting them according to their eigenvalue. At the same time, admitting fluctuations of arbitrarily large amplitude implies that we can still perform the functional integral over all metric configurations, as long as the topology is kept fixed. While quantum Einstein gravity admits a linear split of the metric, unimodularity requires us to use a non-linear split of the form 1 1 g =g¯ exp(h.)κ =g¯ +g¯ hκ+ g¯ hκλh +...=g¯ +h + h hκ+..., (11) µν µκ . ν µν µκ ν 2 µκ λν µν µν 2 µκ ν where the background metric g¯ is used to lower and raise indices at each order of the expansion in the fluctuation µν field h . We then take the path-integral over the fluctuation field h as the definition of the generating functional µν µν for quantum gravity. (The effect of using such a decomposition in quantum Einstein gravity has been studied in in [37, 39] and [38].) Note that a similar decomposition has to be invoked for calculations involving fluctuations, e.g., in a cosmological setting, in the context of (semi-) classical unimodular gravity. We then have at our disposal the background covariant Laplacian −D¯2 which we can use to set up a regulator Rµνκλ(−D¯2+αR¯ (not to be confused with the Riemann tensor) for the fluctuation field h Rµνκλ(−D¯2+αR¯)h , (12) µν κλ where α is the prefactor of a possible additional dependence on the background curvature R¯. While the action is invariant under a simultaneous transformation of the background and fluctuation piece (g¯ → g¯ exp(γ.)λ, h → µκ µλ . κ µκ h −γ +1/2[h ,γ ] −1/12[h ,[h ,γ ]] +1/6[γ ,[γ ,h ]] +...), the regulator term is clearly not. The same µκ µκ .. .. µκ .. .. .. µκ .. .. .. µκ will be true for the gauge-fixing, as we will employ a background gauge fixing here. This implies that background- field couplings and fluctuation-field couplings will satisfy different flow equations. For instance, the RG flow of the background Newton coupling will differ from that of the prefactor of the term quadratic in the fluctuation field and in derivatives. In this work, we will not resolve this difference, but instead identify background and fluctuation-field couplings, and leave the next step to future work. Thus, while a full calculation would feature beta functions for the background and fluctuation couplings where the nontrivial terms can only depend on the fluctuation couplings, our approximation will involve a nontrivial dependence on the background couplings. 2. Second variation of the action Our truncation is given by (cid:90) Γ = d4x(cid:15)f(R). (13) k Within quantum Einstein gravity, an analogous truncation has been considered in [20–24]. TheunimodularityconditionimpliesthatthenumberofpossibletermsinthevariationofEq.(13)willbereduced, √ √ sincetermswhicharepresentinquantumEinsteingravity, suchas(δ2 g)f(R)and(δ g)f(cid:48)(R)δR, donotexisthere. In fact (cid:90) δ2Γ = d4x(cid:15)(cid:0)f(cid:48)(R)δ2R+f(cid:48)(cid:48)(R)(δR)2(cid:1). (14) k Evaluating the variation, starting from the relation Eq. (11), and using a 4-sphere for the background field configu- ration2, we obtain 1(cid:90) (cid:104) (cid:18) 1 1 (cid:19)(cid:105) Γ2 = d4x(cid:15) f(cid:48)(cid:48)(R)h D¯µD¯νD¯κD¯λh +f(cid:48)(R) − R¯h hµν −h D¯µD¯λhν + h D¯2hµν . (15) 2 µν κλ 12 µν µν λ 2 µν Herein, g =g¯ , i.e., we employ a single-metric approximation from now on. µν µν 2 Forthisconfiguration,wehaveR¯µν = R4¯g¯µν andR¯µνκλ= 1R¯2(cid:0)g¯µκg¯νλ−g¯µλg¯νκ(cid:1). 6 Asanextstep,weinsertaYork-decompositionofthefluctuationfieldintoatransversetracelesstensor,atransverse vector, and a scalar (corresponding to the longitudinal vector mode). Note that in contrast to the usual case, there is no trace mode, i.e., 1 h =hTT +D¯ v +D¯ v +D¯ D¯ σ− g¯ D¯2σ, (16) µν µν µ ν ν µ µ ν 4 µν where D¯νhTT =0, g¯µνhTT =0 and D¯µv =0. µν µν µ It turns out that the second variation evaluated on the transverse vector mode vanishes. In other words, the dynamicsofthe vectormodeisarisingfrom thegauge-fixingtermonly, i.e., itis“puregauge”. Thisis anothermajor difference to the case of f(R) truncations in quantum Einstein gravity, see, e.g., [20–22]. For the transverse traceless tensor mode we obtain 1(cid:90) 1(cid:90) (cid:18)1 1 (cid:19) d4x(cid:15)h Γ(2)µνκλh = d4x(cid:15)f(cid:48)(R)hTT D¯2−R¯ hTTµν. (17) 2 µν TT κλ 2 µν 2 12 Finally, the scalar mode is governed by the following dynamics 1(cid:90) 1(cid:90) (cid:104) (cid:18)−1 3 (cid:19) d4x(cid:15)σΓ(2)σ = d4x(cid:15)σ f(cid:48)(R¯) R¯D¯2D¯2− D¯2D¯2D¯2 2 σσ 2 16 16 (cid:18) 9 3 1 (cid:19)(cid:105) +f(cid:48)(cid:48)(R¯) D¯2D¯2D¯2D¯2+ R¯D¯2D¯2D¯2+ R¯2D¯2D¯2 σ. (18) 16 8 16 As usual, no mixed contributions Γ etc. can exist because of the transversality and tracelessness of hTT and v . σv µν µ 3. Gauge-fixing We choose a gauge-fixing that is related to the harmonic gauge condition, but modified such that it satisfies g¯µνD¯ F =0, (19) ν µ for the spherical background. Accordingly, this choice of gauge fixing only imposes three instead of four gauge- conditions, i.e., it only fixes the transversal diffeomorphisms, infinitesimally defined by δ g =L g with D vµ =0. (20) D µν v µν µ Note that in models of gravity which are invariant under transverse diffeomorphism, an additional scalar mode appears upon linearization. As noted in [40], this mode is absent in two cases: If the symmetry is enhanced to full diffeomorphism symmetry, yielding standard Einstein gravity, or if the metric determinant remains fixed. Then the additional scalar, which plays the role of the determinant, is removed from the model. Gauge-fixing only the transverse diffeomorphisms is achieved by using the longitudinal and transversal projectors defined in [41], which read Π = −D¯ (cid:0)−D¯2(cid:1)−1D¯ , (21) Lµν µ ν Π = g¯ −Π . (22) Tµν µν Lµν As they should, these satisfy Π Πν = Π , Π Πν = 0 and Π Πν = Π . We now project the Lµν Lκ Lµκ Lµν Tκ Tµν Tκ Tµκ harmonic gauge on its transversal part [42] and define √ F = 2Πκ D¯νh . (23) µ Tµ νκ Itis thenstraightforward toseethat g¯µνD¯ F =0. Accordingly onlythreeconditions areimposedon the fluctuation ν µ field,whichonecaneasilyseebyinsertingtheYorkdecomposition: Itturnsoutthatthegaugefixingdoesnot impose a condition on σ, but only on v , which has only three independent components. These turn out to be gauge modes. µ Indeed √ (cid:18) R¯(cid:19) F = 2 D¯2+ v . (24) µ 4 µ 7 Thus the gauge-fixing action reads 1 (cid:90) (cid:18) R¯(cid:19)2 S = d4x(cid:15)v D¯2+ vµ. (25) gfv α µ 4 Finally, the Faddeev-Popov ghost action is obtained in the usual way and reads (cid:90) (cid:18) R¯(cid:19) S =− d4x(cid:15)c¯µ D¯2+ c , (26) gh 4 µ where we have already identified g =g¯ and D¯ cµ =0=D¯ c¯µ. (As we only evaluate the ghost loop contribution µν µν µ µ to the running in the gravitational background couplings, this is already allowed at this stage.) 4. Jacobian and auxiliary fields The York decomposition implies the existence of a nontrivial Jacobian in the generating functional [27]. Here, we will deal with this Jacobian by employing the following strategy: From the structure of S it is obvious that a part gfv of the Jacobian can be cancelled by the field redefinition (cid:114) R¯ v → −D¯2− v . (27) µ 4 µ (cid:16) (cid:17) Employing this field redefinition results in S = −1/α(cid:82) d4x(cid:15)v D¯2+ R¯ vµ., i.e., the vector mode does not gfv µ 4 contribute to the RG flow if we impose Landau gauge. In principle, we could choose to nevertheless impose a regulator on that mode with a dependence on the gauge parameter. Here, we take the point of view that a vanishing (unregularized) propagator allows us to trivially integrate out the v mode in the path-integral, such that it does not affect the effective action. (Alternatively, a gauge-choice of the form v = 0 could also be imposed, as in [38], also µ leading to a vanishing contribution of the vector mode.) On the other hand, a corresponding redefinition of σ in order to absorb the remaining part of the Jacobian would not lead to a simple form of the inverse propagator. Accordingly we introduce auxiliary fields to take into account that part of the Jacobian. The corresponding action is given by (cid:90) (cid:104)3 (cid:18) R¯(cid:19) 3 (cid:18) R¯(cid:19) (cid:105) Γ = d4x(cid:15) χ¯ −D¯2− (−D¯2)χ+ ζ −D¯2− (−D¯2)ζ , (28) kaux 4 3 4 3 where χ is a complex Grassmann field and ζ is a real scalar field. These give the same contribution to the flow equation, with a relative factor of −2. Accordingly, the total contribution comes with a factor −1/2. 5. Choice of two regularization schemes and evaluation of traces We will study two different regulators in the following. For the first choice, we follow [22] and employ regulators which essentially substitute the following covariant Laplace-type operators by k2: 2R¯ R¯ R¯ ∆ = ∆+ , ∆ =∆− , ∆ =∆− , (29) 2 12 1 4 0 3 for the transverse traceless tensor, the transverse vector, and the scalar. Herein −D¯2 =∆. We choose a Litim-type cutoff [33] 1 R =− f(cid:48)(R)(k2−∆ )θ(k2−∆ )), (30) kTT1 2 2 2 for the transverse traceless tensor. Note that the negative sign is exactly as it should be, as in the simplest case f(R)= −1 R. For the scalar mode we obtain the slightly lengthier expression 16πG (cid:104) (cid:18)R2 3 9 (cid:19) R = f(cid:48)(cid:48)(R) (k4−∆2)+ R(k6−∆3)+ (k8−∆4) kσ1 16 0 8 0 16 0 (cid:18) 1 1 3 (cid:19)(cid:105) +f(cid:48)(R) R2(k2−∆ )+ R(k4−∆2)+ (k6−∆3) θ(k2−∆ ), (31) 48 0 8 0 16 0 0 8 such that for k2−∆ >0 0 (cid:18)R2k4 3 9 (cid:19) (cid:18) 1 1 3 (cid:19) Γ(2) +R →f(cid:48)(cid:48)(R) + Rk6+ k8 +f(cid:48)(R) R2k2+ Rk4+ k6 . (32) kσ kσ1 16 8 16 48 8 16 Since the function f(R) appears explicitly, its scale-derivatives will feature on the right-hand side of the Wetterich equation. They will lead to a schematic structure of the form ∂ g = c g2+c ∂ g+... of the flow equation for the t 1 2 t couplings. This results in nonperturbative resummation structures, i.e., ∂ g ∼ c1g2. As the prefactor c can contain t 1−c2 2 further couplings, this choice of ”spectrally adjusted” [43, 44] regulator roughly corresponds to a resummation of an infinite series of polynomial terms in the couplings. Explicitly, we will use the following derivatives (cid:16) (cid:17) ∂ f(cid:48)(R) = k2 2f˜(cid:48)+∂ f˜(cid:48)−2R˜f˜(cid:48)(cid:48) , (33) t t ∂ f(cid:48)(cid:48)(R) = ∂ f˜(cid:48)(cid:48)−2R˜f˜(cid:48)(cid:48)(cid:48). (34) t t in terms of the dimensionless function f˜(R˜)=k−4f(R), where R˜ = R. k2 For the auxiliary fields and ghost we take √ R = − 2(k2−∆ )θ(k2−∆ ), (35) kgh1 1 1 (cid:18) (cid:19) R = 3 k4−∆2+ R(cid:0)k2−∆ (cid:1) θ(k2−∆ ). (36) kaux1 4 0 3 0 0 In order to test the reliability of our results, we will actually perform a fixed-point search with two different regu- larization schemes. As our second choice we employ a regulator, which essentially substitutes covariant Laplacians ∆ = −D¯2 with k2 in the regularized propagator. The additional curvature dependence which is introduced in the regulator when using the operators Eq. (29) is removed in this choice, resulting in a shift in possible poles of the flow equation. With ∆=−D¯2, this choice corresponds to 1 R =− f(cid:48)(R)(k2−∆)θ(k2−∆)), (37) kTT2 2 for the transverse traceless tensor. For the scalar mode, we choose a regulator of the form (cid:104) (cid:18)−1 3 (cid:19) R = f(cid:48)(R) R(k4−∆2)+ (k6−∆3) kσ,2 16 16 (cid:18) 9 3 R2 (cid:19)(cid:105) +f(cid:48)(cid:48)(R) (k8−∆4)− R(k6−∆3)+ (k4−∆2) θ(k2−∆). (38) 16 8 16 For the ghost, we choose √ R = 2(−k2+∆)θ(k2−∆). (39) kgh,2 The auxiliary fields come with a regulator of the form (cid:18) (cid:19) 3 R R =− −k4+∆2+ (k2−∆) θ(k2−∆). (40) kaux,2 4 3 Inbothregularizationschemes,wesumovertheeigenvaluesofthecorrespondingLaplacians,whichcanbeobtained forboth∆and∆ fromthefollowingtable,wherethemultiplicitiesarenotaffectedbythecurvature-dependentshift s between ∆ and ∆ . s To convert the sum over eigenvalues into an integral, we employ the Euler-MacLaurin formula. In this step, additional terms which depend on the derivatives of the integrand at the lower boundary, arise. No contributions from the upper boundary exist, as θ(k2 −x) = 0 for x → ∞. From the lower boundary, only the first few terms contribute: As ∂ R (Γ(2) + R ) is a polynomial of finite order in the eigenvalues, only the lowest few orders in t k k k derivatives, when evaluated at the lower boundary, can contribute. 9 eigenvalue multiplicity ∆ n(n+3)−4R; n=0,1,... (n+2)!(2n+3) 0 12 6n! ∆ n(n+3)−4R; n=1,2,... (n+1)!n(n+3)(2n+3) 1 12 2(n+1)! ∆ n(n+3)R; n=2,3,... 5(n+1)!(n+4)(n−1)(2n+3) 2 12 6(n+1)! TABLE I: Eigenvalues and multiplicities of Laplacians acting on transverse traceless tensors, transverse vectors, and scalars, [45]. IV. RESULTS: ASYMPTOTIC SAFETY IN f(R) A. Flow equations The flow equation for f˜(R˜) reads R˜2 ∂ f˜(R˜) = −4f˜(R˜)+2R˜f˜(cid:48)(R˜)+ (F +F +F +F ), (41) t 384π2 TT σ gh aux where the contributions from transverse traceless tensors F , scalars F , auxiliary fields F and Faddeev-Popov TT σ aux ghosts F depend on the choice of regulator, as discussed above. gh For the first choice of cutoff, which is a function of ∆ , we obtain s (cid:18)89 60 40(cid:19) 1 (cid:32)89 20 20 127R˜(cid:33)(cid:16) (cid:17) F = + − + + − − 2f˜(cid:48)+∂ f˜(cid:48)−2f˜(cid:48)(cid:48)R˜ , (42) TT 18 R˜2 R˜ 2f˜(cid:48) 18 R˜2 R˜ 567 t 1 F = · σ (cid:16) (cid:17) 1f˜(cid:48)(cid:48)(3+R˜)2+ 1 f˜(cid:48)(3+R˜)2 8 24 (cid:104)(cid:18) 271 12 (cid:19) (cid:32) (cid:18)9 1 1 (cid:19) (cid:32)9 9 R˜2(cid:33)(cid:33) · − + (1+R˜) · f˜(cid:48) + R˜+ R˜2 +f˜(cid:48)(cid:48) + R˜+ 90 R˜2 8 2 24 2 4 4 (cid:32) (cid:33) (cid:16) (cid:17) 29 27 39 3 12161R˜3 + 2f˜(cid:48)+∂ f˜(cid:48)−2f˜(cid:48)(cid:48)R˜ − + + −R˜− R˜2+ t 48 20R˜2 16R˜ 16 6842880 (cid:32) (cid:33) (cid:16) (cid:17) 21 9 81 23R˜ 9R˜2 439573R˜4 (cid:105) + ∂ f˜(cid:48)(cid:48)−2f˜(cid:48)(cid:48)(cid:48)R˜ − + + − − − , (43) t 16 2R˜2 10R˜ 8 16 622702080 (cid:32) (cid:33) 109 36+24R˜ F = −2 + , (44) gh 30 R˜2 6+R˜ (cid:16) (cid:17) F = −1080+R˜(−1080+271R˜) . (45) aux 90R˜2(3+R˜) As a main structural difference to quantum Einstein gravity with a full dynamical metric, including a conformal factor [20–22], one should note that the function f˜does not appear on the right-hand side of the flow equation, i.e., the cosmological constant is not part of the theory space. 10 For the second regulator, that imposes cutoffs on ∆, we obtain 1 (cid:104) (cid:16) (cid:17) F = −252f˜(cid:48) −1080+R˜(360+R˜) TT (cid:16) (cid:17) 4536R˜2f˜(cid:48) 1+ R˜ 6 (cid:16) (cid:17)(cid:16) (cid:16) (cid:17)(cid:17)(cid:105) + 2f˜(cid:48)+∂ f˜(cid:48)−2f˜(cid:48)(cid:48)R˜ 45360R˜+R˜ −22680+R˜(−126+311R˜) (46) t 1 F = · σ (cid:16) (cid:17) 2 3−R˜f˜(cid:48)+f˜(cid:48)(cid:48)9−6R˜+R˜2 16 16 (cid:34) (cid:32) (cid:33) 1 631 72 72 551R˜ 511R˜2 55189R˜4 · (∂ f˜(cid:48)(cid:48)−2R˜f˜(cid:48)(cid:48)(cid:48)) − + − + − − 16 t 10 R˜2 5R˜ 15 90 38918880 1 (cid:16) (cid:17)(cid:16) (cid:17) + 2f˜(cid:48)+∂ f˜(cid:48)−2f˜(cid:48)(cid:48)R˜ 64665216+8981280R˜−58977072R˜2+16997904R˜3+3815R˜5 47900160R˜2 t (cid:32) (cid:32) (cid:33) (cid:32) (cid:33)(cid:33) (cid:32) (cid:33)(cid:35) 9 R˜ 9 9 R˜2 511 4(3+R˜) + f˜(cid:48) − +f˜(cid:48)(cid:48) − R˜+ · − + (47) 8 4 2 4 4 90 R˜2 (cid:32) (cid:33) 8 7 6(6+R˜) F = − − + , (48) gh 4−R˜ 60 R˜2 −6+R˜ (cid:16) (cid:17) F = −1080+R˜(−360+511R˜) . (49) aux 90R˜2(−3+R˜) Comparing the two equations for ∂ f˜(R˜), we note that the different choice of regularization scheme mainly serves t the purpose of changing the singularity structure of the equation, which can play a role in the search for global solutions[22]. Apartfromthat,thestructureissimilarinbothcases,withdifferencesonlyinthenumericalprefactors of most terms. B. Fixed points To search for fixed points, we expand f˜(R˜) in a polynomial around vanishing curvature, where the cosmological constant is again absent, 10 f˜(R˜)= (cid:88)a R˜n. (50) n n=1 Note that the Newton coupling is given by G = − 1 . Thus a negative fixed-point value for a translates into a 16πa1 1 positive microscopic Newton coupling. Using the first regularization choice Eq. (42) - Eq. (45), we find a number of fixed points, where we list only the most stable one in Tab. II. Beyond R˜4 it becomes rather cumbersome to evaluate all solutions of the fixed-point equations. We thus pick the most stable fixed point that exists in the four truncations up to a . At R˜5 and beyond we perform a numerical search for the solution of the fixed-point equation in the vicinity 4 of the fixed-point coordinates at lower truncation order, and no longer investigate all solutions. The fixed point at n = 1 is related to that found in [29], which neglected the additional terms arising from the Euler MacLaurin formula. We clearly observe that the fixed point at truncation order n+1 is a continuation of the fixed point at order n, as the coordinates of the fixed points lie reasonably close to each other. This stability under extensions of the truncation indicates that this is not an auxiliary fixed point. Interestingly, the fixed-point values of the couplings a , n ≥ 5 are four orders of magnitude smaller than the highest-order coupling. This suggests that n an approximation of the fixed-point action employing only the first few terms could be reasonable, for instance if cosmological consequences of unimodular asymptotic safety are deduced from ”RG improved” calculations. The corresponding critical exponents are given in Tab. III, and again seem reasonably stable under extensions of the truncation, cf. Fig. 1. Similarly to the case of quantum Einstein gravity, two exponents with positive real part are observed, i.e., R and R2 form relevant directions. In this approximation, the two critical exponents with positive real part form a complex conjugate pair, which could hint at a necessary extension of the truncation to obtain better numerical precision. Higher-order operators become increasingly irrelevant. The anomalous dimension is a positive contributiontothecriticalexponentsforalloperators,i.e.,quantumfluctuationsshiftalloperatorstowardsrelevance. Comparing the value of the critical exponents θ beyond i=3 to the canonical dimension d =−(2i−4), we observe i i