Zooming in on fermions and quantum gravity
Abstract
We zoom in on the microscopic dynamics for fermions and quantum gravity within the asymptoticsafety paradigm. A key finding of our study is the unavoidable presence of a nonminimal derivative coupling between the curvature and fermion fields in the ultraviolet. Its backreaction on the properties of the Reuter fixed point remains small for finite fermion numbers within a bounded range. This constitutes a nontrivial test of the asymptoticsafety scenario for gravity and fermionic matter, additionally supplemented by our studies of the momentumdependent vertex flow which indicate the subleading nature of higherderivative couplings. Moreover our study provides further indications that the critical surface of the Reuter fixed point has a low dimensionality even in the presence of matter.
I Introduction
In the search for a quantum theory of gravity that is viable in our universe, the existence of fermionic matter must be accounted for. Our strategy to achieve this is based on a quantum field theoretic framework that includes the metric field and fermion fields at the microscopic level. Such a setting requires an ultraviolet completion or extension of the effective field theory framework within which a joint description of gravity and matter is possible up to energies close to the Planck scale. Asymptotic safety Weinberg (1979); Reuter (1998) is the idea that scaleinvariance provides a way to extend the dynamics to arbitrarily high momentum scales without running into Landau poles which would indicate a triviality problem. Moreover, scaleinvariance is a powerful dynamical principle, that is expected to fix all but a finite number of free parameters in an infinite dimensional space of theories. It can be reached at a fixed point of the Renormalization Group (RG), which can be free (asymptotic freedom) or interacting (asymptotic safety). Compelling indications for the existence of the asymptotically safe Reuter fixed point in fourdimensional gravity have been found, e.g., in Reuter and Saueressig (2002); Lauscher and Reuter (2001); Litim (2004); Codello et al. (2009); Benedetti et al. (2009); Manrique et al. (2011); Becker and Reuter (2014); Demmel et al. (2015); Gies et al. (2016); Denz et al. (2018); Christiansen et al. (2018a); Falls et al. (2018a). For recent reviews and introductions including a discussion of open questions, see Reuter and Saueressig (2012); Ashtekar et al. (2014); Eichhorn (2018a); Percacci (2017); Eichhorn (2018b).
A central part of the interplay of the Standard Model with gravity is the impact of quantum gravity on the microscopic dynamics for fermions as well as the corresponding “backreaction” of fermionic matter on the quantum structure of spacetime. In line with the observation that asymptotically safe quantum gravity could be nearperturbative Eichhorn et al. (2018a, b) or “as Gaussian as it gets” Falls et al. (2013, 2016, 2018b, 2018a), studies of fermiongravitysystems follow a truncation scheme by canonical power counting. Furthermore, the chiral structure of the fermion sector of the Standard Model is a key guiding principle. Thus, the leadingorder terms according to canonical power counting have been explored in the sector of chirally symmetric fermion self interactions Eichhorn and Gies (2011); Meibohm and Pawlowski (2016); Eichhorn and Held (2017), and fermionscalar interaction sector Eichhorn et al. (2016a); Eichhorn and Held (2017). These are dimension6 and dimension8operators, respectively. Explicitly chiralsymmetry breaking interactions, including a mass term and two dimension5, nonminimal couplings of fermions to gravity Eichhorn and Lippoldt (2017), have been studied. The effect of quantum gravity on a Yukawa coupling of fermions to scalars has been studied in Zanusso et al. (2010); Vacca and Zanusso (2010); Oda and Yamada (2016); Eichhorn et al. (2016a); Hamada and Yamada (2017); Eichhorn and Held (2017, 2018a). Conversely, the impact of fermionic fluctuations on the Reuter fixed point has been explored in Donà and Percacci (2013); Donà et al. (2014); Meibohm et al. (2016); Eichhorn and Lippoldt (2017); Biemans et al. (2017a); Alkofer and Saueressig (2018); Alkofer (2019). An asymptotically safe fixed point exists in all of these studies, as long as the fermion number is sufficiently small. Moreover, all operators that have been explored follow the pattern that canonical dimensionality is a robust predictor of relevance at the interacting fixed point. Further, they confirm the conjecture that asymptotically safe quantum gravity could preserve global symmetries Eichhorn and Held (2017), at least in the Euclidean regime. Thus, all symmetrybreaking interactions can be set to zero consistently. Additionally, interacting fixed points could, but need not exist for these, as in the case of the Yukawa coupling Zanusso et al. (2010); Vacca and Zanusso (2010); Oda and Yamada (2016); Eichhorn et al. (2016a); Hamada and Yamada (2017); Eichhorn and Held (2017, 2018a). In contrast, the interacting nature of asymptotically safe gravity percolates into the symmetric sector, where interactions can typically not be set to zero consistently Eichhorn (2012); Eichhorn and Held (2017). Hence, their “backreaction” on the asymptotically safe fixed point could be critical. Further, this sector is a potential source of important constraints on the microscopic gravitational parameter space: Strong gravity fluctuations could trigger new divergences in the matter sector, manifesting themselves in complex fixedpoint values for matter interactions. The corresponding bound on the gravitational parameter space that separates the allowed, weakly coupled gravity regime from the forbidden stronglycoupled regime, is called the weakgravity bound Eichhorn et al. (2016a); Christiansen and Eichhorn (2017); Eichhorn and Held (2017).
In Tab. 1 we provide an overview over interactions in the fermion sector that have been explored in an asymptotically safe context. The table contains a crucial gap, namely nonminimal, chirally symmetric interactions.
ref.  interaction  dimension  relevant  symmetry  free fixed point  weakgravity bound 

Eichhorn and Lippoldt (2017) 

3  yes  no sym.  yes  no 
Eichhorn and Held (2017) 

4  both possible  no sym.  yes  no 
Eichhorn and Lippoldt (2017) 

5  no  no sym.  yes  no 
Eichhorn and Lippoldt (2017) 

5  no  no sym.  yes  no 
this work 

6  no  sym.  no  no 
Eichhorn and Gies (2011); Meibohm and Pawlowski (2016); Eichhorn and Held (2017) 

6  no  sym.  no  no 
Eichhorn and Gies (2011); Meibohm and Pawlowski (2016); Eichhorn and Held (2017)] 

6  no  sym.  yes  no 
Eichhorn et al. (2016a); Eichhorn and Held (2017) 
, 
8  no  sym.  no  yes 
This is the sector that we will begin to tackle in this paper. For an analogous study in the scalar sector see Eichhorn et al. (2018c).
As our key result we find a continuation of the asymptotically safe Reuter fixed point to finite fermion numbers that passes a nontrivial test by remaining robust under a crucial extension of the approximation to the full dynamics. Moreover, we find further indications that the critical hypersurface of the Reuter fixed point has a low dimensionality also in the presence of matter.
This paper is structured as follows: In Sec. II, we provide an overview of the setup, and specify the approximation to the full dynamics that we will explore in the following. In Sec. III we discuss in some detail how to derive the beta functions in our setting. In particular, we discuss the relation of the derivative expansion to the projection at finite momenta. Sec. IV provides an overview of the fixedpoint results for , which are representative for the results at small fermion numbers. We discuss tests of the robustness of the fixed point, the impact of the newly included nonminimal derivative interaction on the fixedpoint results in a smaller truncation, and the feature of effective universality. Sec. V contains a discussion of structural aspects of the weakgravity bound for cubic beta functions and highlights that no such bound exists for the nonminimal derivative interaction in the regime of gravitational parameter space where our truncation remains viable. In Sec. VI we extend our investigations to , and discuss the continuation of the Reuter fixed point to larger fermion numbers. In Sec. VII we provide a short summary of our key results and highlight possible routes forward in gravitymatter systems in an outlook. App. A includes a general derivation of the form of the flow equation for the dimensionless effective action. This form can be used to directly derive dimensionless beta functions, in contrast to the usual procedure of only introducing dimensionless quantities after a truncation has been specified.
Ii Setup
The system we analyze contains a gravitational sector and a matter sector with chiral fermions. We aim at deriving the beta functions in this system, and will employ the wellsuited functional Renormalization Group. It is based on the flow equation for the scaledependent effective action, the Wetterichequation Wetterich (1993); Ellwanger (1994); Morris (1994),
(1) 
The “superfield” is simply a collection of all fields in our system,
(2) 
where Einsteins summation convention over the “superindex” contains a summation over discrete spacetime, spinor and flavor indices and an integration over the continuous coordinates. Here is a scaledependent regulator that implements a momentumshell wise integration of quantum fluctuations and the dot in refers to a derivative with respect to , the RG“time” with an arbitrary reference scale. The IRregulator enters the generating functional in the form of a term that is quadratic in the fluctuation fields and renders the Wetterich equation UV and IR finite. Specifically, we choose a Litimtype cutoff Litim (2001) with appropriate factors of the wavefunction renormalization for all fields. Next to the gaugefixing term for the metric fluctuations, it is a second source of breaking of diffeomorphism invariance. It must be set up with respect to an auxiliary metric background , which provides a notion of locality and thereby enables a local form of coarse graining. In the main part of this paper we focus on a flat background,
(3) 
while in this section we will keep arbitrary for pedagogical reasons. For introductions and reviews of the method, see, e.g., Berges et al. (2002); Delamotte (2007); Rosten (2012); Braun (2012); specifically for gauge theories and gravity, see, e.g., Pawlowski (2007); Gies (2012); Reuter and Saueressig (2012).
The Wetterich equation provides a tower of coupled differential equations for the scale dependence of all infinitely many couplings in theory space. In practice, this has to be truncated to a (typically) finitedimensional tower. Let us briefly summarize how we proceed, before providing more details. To construct our truncation, we define a diffeomorphism invariant “seed action”. Next, we expand the terms in this seed action to fifth order in metric fluctuations, defined as
(4) 
This corresponds to an expansion of the seed action in vertices. At this point all terms in the seed action, except those arising from the kinetic term for fermions, come with one of the couplings of the seed action. We next take into account that in the presence of a regulator and gauge fixing, the beta functions for those couplings generically differ, when extracted from different terms. Accordingly, we introduce a separate coupling in front of each term in the expanded action. This provides the truncation which we analyze in the following. To close the truncation, the couplings of higherorder vertices are partially identified with those of lowerorder ones.
In more detail, these steps take the following form: Our seed action reads
(5) 
Classical gravity is described by the EinsteinHilbert action ,
(6) 
In order to tame the diffeomorphism symmetry of gravity, we choose a gaugefixing condition ,
(7) 
The gauge choice is incorporated using the gauge fixing action ,
(8) 
To take care of the resulting FaddeevPopov determinant, we use ghost fields and with the appropriate ghost action ,
(9) 
where is the Lie derivative of the full metric in ghost direction,
(10) 
In the following, we choose the Landau gauge, i.e., . By employing a York decomposition of we see that this choice of gaugefixing parameters leads to contributions from only a transversetraceless (TT) mode and a trace mode ,
(11) 
where the TTmode satisfies and , while the trace mode is given by . All other modes drop out of the flow equation once it is projected onto monomials with nonvanishing powers of the field. It is important to note that the TTmode is present in any gauge and to linear order in a gauge invariant quantity. Thus, for external metric fluctuations we exclusively consider the TTmode. For internal metric fluctuations, also the remaining trace mode is taken into account. We summarize the purely gravitational parts of the action as ,
(12) 
Next we turn to the chiral fermions. Their minimal coupling to gravity is via the kinetic term ,
(13) 
For the construction of the covariant derivative for fermions, we use the spinbase invariance formalism Gies and Lippoldt (2014, 2015); Lippoldt (2015). For our purposes, this is equivalent to using the vierbein formalism with a Lorentz symmetric gauge. Upon expansion in , this minimal interaction between fermions and gravity gives rise to an invariant linear in derivatives. There are several invariants containing terms of third order in derivatives and canonical mass dimension, namely:
where each of the invariants respects the OsterwalderSchrader positivity of the Euclidean action. Out of these four invariants, the ones corresponding to and do not contribute linearly to an external , as does not contain a transverse traceless part to linear order. In the following, we restrict ourselves to the nonminimal coupling and neglect the term. Thus, the kinetic matter action is complemented with
(15) 
has all the symmetries of the original action (5) and therefore does not enlarge the theory space.
The nonminimal coupling introduces an invariant of cubic order in derivatives, capturing parts of the higherderivative structure of the fermiongravity interaction. Once expanded around a flat background, the interaction with is given by
(16) 
where is the d’Alambertian in flat Euclidean space. Eq. (16) is the unique invariant consisting of one , , and together with two derivatives acting on the TTmode and one derivative acting on the . We summarize the matter parts of the action as ,
(17) 
After having specified our complete seed action, we expand the scaledependent effective action in powers of the fluctuation field,
(18) 
where refers to functional derivatives with respect to the field ,
(19) 
Note the order of the indices and fields, which is important to keep in mind for the Grassmannvalued quantities.
By using this vertex form, the flow of 5 individual couplings , , , and as well as the anomalous dimension of two wavefunction renormalizations and is disentangled, cf. Tab. 2 and see Sect. III for more details.
Couplings  

,  –  
,  
,  
, 
Here the barred couplings, e.g., and , refer to dimensionful couplings.
For the gravityfermion vertex the contributing diagrams are shown in Fig. 1. This highlights the necessity to truncate the tower of vertices, as the flow of each point vertex depends on the  and point vertices. We use the seed action in Eq. (5) to parametrize the vertices appearing in the diagrams. When generating, e.g., a graviton threepoint vertex or a graviton fourpoint vertex for the scaledependent effective action from the seed action by expanding to the appropriate power in , both would depend on the same Newton coupling and the same cosmological constant due to diffeomorphism symmetry. However, the gauge fixing and the regulator break diffeomorphism symmetry. Hence, the effective action is known to satisfy SlavnovTaylor identities instead, Ellwanger et al. (1996); Reuter (1998); Pawlowski (2007, 2003); Manrique and Reuter (2010); Donkin and Pawlowski (2012). As these identities in general are much more involved, there is no such simple relation between the three and fourpoint vertex of the effective action as there is for the seed action. In other words, the breaking of diffeomorphism symmetry leads to an enlargement of theory space in which the couplings parameterizing the vertices are independent. There are different routes towards a truncation of this large theory space. In principle, one could pick some random tensor structure and momentumdependence in each point function and parameterize this by some coupling. Then, the connection to the diffeomorphisminvariant seed action would be lost completely. Instead, we derive the tensor structures of the vertices from the seedaction, but also take into account that the various couplings are now independent. Specifically, we proceed using the following recipe: The structure of the point vertex is drawn from the seed action,
(20) 
where the replacement only affects pure gravity vertices. Furthermore in Eq. (20) the metric fluctuations of the purely gravitational action are rescaled according to
(21) 
whereas the graviton in and is rescaled to
(22)  
(23) 
This rescaling breaks diffeomorphism symmetry and helps us choosing a basis in the appropriate theory space. Note that the fieldredefinitions in Eq. (21), (22) and (23) are not to be understood as actual fieldredefinitions in the effective action. They are just a way of arriving at a parameterization of the truncated effective action in the enlarged theory space.
In the following we use the term “avatar”, when a single coupling in the seed action leads to various incarnations
in the effective action, e.g., and are avatars of the Newton coupling .
In order to close the flow equation, we identify couplings of higher order point vertices with the corresponding couplings of the threepoint vertex. This was already implicitly done with the rescaling in Eqs. (21), (22) and (23) and with the usage of one single coupling . Similarly, all point vertices arising from the cosmologicalconstant part of the seed action are parametrized by one coupling , i.e., and . The relation between and the gravitational massparameter that is often used in the literature, reads . In the next section we provide details on how the beta functions are extracted from the sum of the diagrams in Fig. 1.
Iii How to obtain beta functions
We now discuss in some detail how to derive beta functions. We concentrate on the dimensionless couplings, which are obtained from their dimensionful counterparts by a multiplication with an appropriate power of . Dimensionful couplings are denoted with overbars, e.g., etc., whereas their dimensionless counterparts lack the overbar, e.g., etc.
A key goal of ours is to test the quality of our truncation. Thus, we place a main focus on the momentumdependence of the flow, i.e., the dependence of the point vertices on the momenta of the fields. Higherorder momentumdependencies than those included in the truncation are in general present. This implies that different projection schemes might yield different results when working in truncations. We will discuss these different schemes and their relation to each other in the following.
iii.1 Fermionic Example
As a concrete example let us consider the fermionic sector. To arrive at beta functions, we have to take several steps. First we define a projector on the gravityfermion vertex. Its form is motivated by the tensor structure of the considered threepoint function,
(24)  
By taking the corresponding functional derivatives of Eq. (24) and evaluating in momentum space, while using the projector onto transverse traceless symmetric tensors , we find that,
(25)  
Of the three momenta, only two are independent, the third can be eliminated by momentum conservation. Thus we define the projector on as
(26) 
which we evaluate at the symmetric point for the momenta, . The normalization of follows from
(27) 
Using we define the projected dimensionful vertex as
(28) 
where implies the trace over Dirac and flavor indices. This definition is independent of any truncation, while a truncation for can be viewed as choosing a specific point in theory space. For instance, when evaluating for our chosen truncation we find that is equal to . Having defined , we aim at deriving the beta function for the dimensionless counterpart ,
(29) 
Note that carries a nontrivial dimension, as the gravityfermion vertex contains an additional momentum . The scale derivative of reads
(30) 
where one has to take into account the scaling of the momentum, . Here and are the anomalous dimensions,
(31) 
We can read off by replacing with in equation (28),
(32) 
where is a short hand for the contributing diagrams in Fig. 1,
(33) 
Here we made use of Eq. (1).
By inserting the expression for given in Eq. (32) into Eq. (30) for we finally arrive at the beta function for ,
(34)  
This equation will take center stage in our analysis of the momentumdependence of the flow and tests of robustness of the truncation. We highlight that in general the righthandside of the flow equation generates terms beyond the chosen truncation. In Eq. (34) the consequence is, that our truncation does not capture the full momentumdependence that is generated. Accordingly, the fixedpoint equation cannot be satisfied for all momenta, but instead only at selected points. We will extensively test how large the deviations of from zero are in order to judge the quality of different truncations.
iii.2 Projection schemes
We perform our analysis in several different projection schemes, as a comparison between the fixedpoint structure of the different truncations provides indications for or against the robustness of the fixed point. We now motivate the use and explain the details of these three projection schemes.
Using Eq. (34), the momentumdependent fixedpoint vertex could be found by demanding and solving Eq. (34). In practice, we choose an ansatz for , which is part of choosing a truncation. At a point in theory space defined by , Eq. (34) holds, but indicates that terms not yet captured by are generated. These are present in Eq. (34), so that we need to truncate the beta function in order to close the system. For example, in our setup, we restrict to a polynomial up to first order in , i.e.,
(35) 
and
(36) 
Here we introduced a modified version of the coupling ,
(37) 
where and are the dimensionless counterparts of and ,
(38) 
However, this specific ansatz does not satisfy Eq. (34) for all values of . Accordingly, the righthand side of Eq. (34) differs from Eq. (36). This is simply an example for the general fact that, plugging a truncation into the righthandside of the Wetterich equation, terms beyond the truncation are generated and therefore the truncation is not closed.
As is not equal to for all momenta, we can choose selected points in the interval for which we demand that is exactly equal to at these points, see, e.g., Eqs. (40) and (41). However, we can also choose superpositions of more values for , see, e.g., Eq. (42) for . Even though this superposition might lead to being not exactly equal to at any point, it can still lead to an overall better description of the full momentum dependence, by being almost equal in a larger region. The values of the coefficients in the ansatz, i.e., and , depend on this choice.
Let us now compare two popular choices, namely the derivative expansion about , and a projection at various values for . Working within a derivative expansion about , one extracts the flow of the th coefficient of the polynomial by the th derivative of Eq. (34), evaluated at . Specifically, for the chosen ansatz in Eq. (35) together with Eq. (36), this yields:
(39) 
This expansion ensures that and its derivative are equal to and its derivative at . However, the derivative expansion to this order does not satisfy this equality away from . This simply means that higherorder terms in the derivative expansion around are generated by the flow. By the evaluation at a single point in , this scheme is very sensitive to local fluctuations at , which might cause deviations for larger momenta.
Alternatively, we can choose finite momenta, e.g., , to extract one of the beta functions. Equating and at and , and solving for the beta functions yields
(40)  
(41) 
In this scheme, the beta functions and by construction are equal at and . Thus, it provides an interpolation between these momenta, while the derivative expansion provides an extrapolation from onwards. The same projection schemes can analogously be applied to other point functions, including the anomalous dimensions. We will refer to the projection at different values for the momentum by samplepoint projection in the following. More specifically, starting from Eq. (40), the beta functions for and take the following form
(42)  
(43) 
with
(44) 
In practice, the ingredients to evaluate the beta functions are the following: is given by the sum of diagrams in Fig. 1 which uses xAct Brizuela et al. (2009, 2009); MartínGarcía (2008); MartínGarcía et al. (2007, 2008) as well as the FORMtracer Cyrol et al. (2017), is given by Eq. (36) and the anomalous dimensions are extracted from a projection of the corresponding twopoint functions at , as in Christiansen et al. (2016, 2015); Meibohm et al. (2016); Denz et al. (2018).
We now provide our motivations for using projections with and sampling points. The derivative expansion at for the gravitymatter avatars of the Newton coupling does not capture all properties of the flow in a quantitatively reliable way cf. the discussion in Eichhorn et al. (2018a). In particular, a derivative expansion of the EinsteinHilbert truncation at , together with a momentumindependent anomalous dimension for the graviton results in a slightly screening property of gravity fluctuations on the Newton coupling. We expect that at higher orders in the truncation, the derivative expansion becomes quantitatively reliable, but in our truncation projections at finite momenta are preferable.
Instead, an expansion at finite momenta is expected to be more stable in small truncations. This is easiest to appreciate when thinking of the flow equation in terms of a vertex expansion: The point functions that enter the flow depend on momenta. Of these, one becomes the loop momentum in the flow equation. Due to the properties of the regulator, the momentum integral over the loop momentum is peaked at . Accordingly, the flow depends on the vertex at a finite momentum, not vanishing momentum, cf. Fig. 2.
Accordingly, a good approximation of the full flow might require higher orders in the derivative expansion around than in projection schemes at finite momentum. For technical simplicity, a symmetric point where the magnitudes of all momenta at the vertex are chosen to be the same nonzero value is preferable, although the example in Fig. 2 showcases that a nonsymmetric point is likely to most accurately capture the momentumdependence of the vertex as it is relevant for the feedback into the flow equation.
We point out that for this type of projections, a onetoone mapping between the couplings extracted in this way and the couplings of the action written in a derivative expansion in terms of curvature invariants, as it is usually done, becomes more involved. For the derivative expansion about , this mapping is onetoone. Specifically, projecting onto a term at finite yields a different result than projection at vanishing . This remains the case even in untruncated theory space, where the couplings in a derivative expansion around zero momentum and the couplings in a projection at finite momenta satisfy a nontrivial mapping onto each other. In an untruncated theory space, such a difference in the choice of basis does not matter for the universal properties of the fixed point. In truncations, such choices can make a difference, as some expansions are better suited to capturing the flow already in small truncations. One might tentatively interpret the results in the EinsteinHilbert truncation and small extensions thereof Christiansen et al. (2016, 2015); Meibohm et al. (2016); Denz et al. (2018); Eichhorn et al. (2018a); Christiansen et al. (2018b); Eichhorn et al. (2018b) as implying that projections with sampling points are preferred over the derivative expansion about vanishing momentum.
For the fermiongravity vertex, we consider the following three approximations
For completeness let us explain how we extract the remaining gravitational couplings, , , , and the wavefunction renormalizations, , . An analogous parameterization to Eqs. (28) and (35) holds for the twofermion, as well as the two and threegraviton vertices. Thus, in all approximation schemes under consideration, we use a projection with samplepoints of and at and for these couplings, i.e., we define them as follows:
(45)  
where the momenta and are evaluated at the symmetric point for three momenta with and the normalizations , , , and are defined such, that when we plug the from Eqs. (20), (21), (22) and (23) into Eq. (45) we get the corresponding coupling.
Iv Asymptotic safety for one flavor
Phenomenologically, fermiongravity systems with are of most interest, as this is the number of Dirac fermions in the Standard Model, extended by three righthanded neutrinos. There are indications Donà and Percacci (2013); Donà et al. (2014); Meibohm et al. (2016); Eichhorn and Lippoldt (2017); Alkofer and Saueressig (2018); Alkofer (2019) that such a fermiongravity system with features an asymptotically safe fixed point that is continuously connected to the puregravity one. We explore this hypothesis further, and therefore start by exploring a small deformation of the puregravity universality class by fermions.
In this section, we aim at answering three key questions:

Is there a fixed point in the fermiongravity system that is robust under extensions of the truncation and changes of the projection scheme?

Is the nonminimal coupling nonzero at the fixed point, and how large is its “backreaction” onto the minimally coupled system?

Do the avatars of the Newton coupling exhibit effective universality at this fixed point?
iv.1 A fermiongravity fixed point and tests of its robustness
Trunc 
