# Relativistic second gradient theory of continuous media

Mina Chapon, Lionel Darondeau, Rodrigue Desmorat, Clément Ecker, Boris Kolev

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Mina Chapon, Lionel Darondeau, Rodrigue Desmorat, Clément Ecker, Boris Kolev. Relativistic second gradient theory of continuous media. 2024. hal-04792877v2

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# RELATIVISTIC SECOND GRADIENT THEORY OF CONTINUOUS MEDIA

M. CHAPON, L. DARONDEAU, R. DESMORAT, C. ECKER, AND B. KOLEV

**ABSTRACT.** Variational Relativity is a framework developed by Souriau in the sixties to better formulate General Relativity and its classical limit: Classical Continuum Mechanics. It has been used, for instance, to formulate Hyperelasticity in General Relativity. In that case, two primary variables are involved, the universe (Lorentzian) metric $g$ and the matter field $\Psi$. A Lagrangian density depending on the 1-jet of these variables is then introduced which must satisfy the principle of General Covariance. Souriau proved in 1958 that under these hypotheses, the Lagrangian density depends only on the punctual value of the matter field $\Psi$ and of a secondary variable $K$, the conformation, an invariant of the diffeomorphism group, which is the Relativistic analog of the inverse of the right Cauchy-Green tensor. In the present work, an extension of Souriau's results to a second order gradient theory in General Relativity is presented. Accordingly, new higher order diffeomorphisms invariants are found. Their classical limits are calculated, showing that the 3-dimensional Continuum Mechanics second gradient theory can be derived from such a relativistic theory. Some of these invariants converge to objective quantities in the Galilean limit, others to non-objective quantities. The present work contributes thus to clarify the theoretical foundation of higher gradient Continuum Mechanics theory.

## CONTENTS

<table><tr><td>1. Introduction</td><td>2</td></tr><tr><td>2. Matter fields in Variational Relativity</td><td>3</td></tr><tr><td>2.1. General Covariance</td><td>4</td></tr><tr><td>2.2. Variational Relativity</td><td>5</td></tr><tr><td>2.3. Matter field</td><td>6</td></tr><tr><td>2.4. Matter rest frame</td><td>7</td></tr><tr><td>3. General covariant Lagrangians involving matter fields</td><td>9</td></tr><tr><td>3.1. First Gradient Relativistic Hyperelasticity</td><td>9</td></tr><tr><td>3.2. Second Gradient Relativistic Hyperelasticity</td><td>9</td></tr><tr><td>3.3. Gradient Fluids</td><td>13</td></tr><tr><td>4. Time, Space and observer frames</td><td>15</td></tr><tr><td>4.1. Observer</td><td>15</td></tr><tr><td>4.2. Rewriting the invariants in the observer frame</td><td>16</td></tr><tr><td>4.3. Second Gradient Hyperelasticity in Minkowski spacetime</td><td>17</td></tr><tr><td>5. Classical limit of Gradient Relativistic Hyperelasticity</td><td>20</td></tr><tr><td>5.1. Classical limit of the relativistic invariants</td><td>20</td></tr><tr><td>5.2. From General Covariance to Objectivity</td><td>21</td></tr><tr><td>6. Conclusion</td><td>21</td></tr><tr><td>Appendix A. Covariant derivatives, Torsion, Curvature and Hessian tensors</td><td>22</td></tr><tr><td>A.1. Covariant derivatives</td><td>22</td></tr><tr><td>A.2. Action of the diffeomorphism group</td><td>23</td></tr><tr><td>A.3. Normalization of the Christoffel symbols</td><td>25</td></tr><tr><td>Appendix B. Volume forms and Hodge operator</td><td>27</td></tr><tr><td>B.1. Volume pseudo-form</td><td>27</td></tr><tr><td>B.2. Hodge star operator</td><td>27</td></tr><tr><td>B.3. Matter coframe</td><td>28</td></tr><tr><td>References</td><td>29</td></tr></table>

Date: July 9, 2025.

2020 Mathematics Subject Classification. 74B20; 83C55; 70G45.

**Key words and phrases.** Relativistic elasticity, Second gradient theory of continuous media, Variational Relativity, General covariance, Objectivity.

Authors names are in alphabetical order.
# 1. INTRODUCTION

Continuum Mechanics has a longstanding tradition of enhancement [47, 20, 46, 21, 44, 4, 23, 57, 64], by incorporating into constitutive equations and conservation laws time derivatives and spatial gradients, i.e., jets of kinematic variables. Fluid Mechanics and Solid Mechanics have however evolved along different trajectories due to the distinct nature of the quantities and settings they address, but also due to the role given to the Principle of Objectivity [52, 68], as fundamental or not for their modeling.

The first aim of the present work is to precise the theoretical foundations of a higher gradient Continuum Mechanics theory:

(1) by deriving second order jet Continuum Mechanics from General Relativity, starting from the Principle of General Covariance,

(2) by formulating Classical (meaning here 3-dimensional and non-relativistic) Second Gradient Hyperelasticity as the Galilean limit ($c \to \infty$) of the relativistic theory when expressed in the Minkowski spacetime [61].

The second aim is to provide an answer, even partial, to the question

> "Is Objectivity (Frame indifference) a consequence of General Covariance ?"

We refer to the work of Germain [28, 29] for an introduction to the (non relativistic) kinematics of higher gradients media. We simply recall that the finite strain gradient theory introduces an internal length in Solid Mechanics by considering in the energy density the additional kinematics variable $F_{\phi}^{-1}\nabla F_{\phi}$, which is objective and built from the spatial second derivative $\nabla F_{\phi} = (\partial^2 x^i / \partial X^J \partial X^K)$ of the deformation $\phi: X \mapsto x = \phi(X)$ (see [46, 28, 22, 24, 14]). An interesting special case corresponds to the small strain formulation by Aifantis and coworkers [1, 2]. Advanced constitutive theories introduce additional higher order stress tensors [13, 67, 26, 25], possibly non-symmetric [66, 47] (dual to the higher order kinematics variables, see also [15, 8, 9] and [64, Supplement 7.2]). Note that purely spatial gradients seem to be absent from fluids constitutive laws. Lastly, the specific kinematic energy $\frac{1}{2}\|\partial_t\phi\|^2$ is half the square norm of the (Lagrangian) velocity $\partial_t\phi$, which is itself a first order time derivative indeed present in the Lagrangian density of deformed media. But higher order time derivatives $\partial_t^k\phi$ of the deformation $\phi$ are usually not considered in Classical Continuum Mechanics Lagrangian densities (even when micro-inertia is accounted for [45]).

Elasticity has been formulated in Special Relativity [32, 42, 55, 56, 51] and, since 1918, in General Relativity at the astrophysics scale [53, 59, 65, 16, 58, 7, 12, 35, 6, 19, 11]. Introducing the framework of Variational Relativity [59, 60, 61] (adopted in [12, 35, 36, 6]), Souriau was able to model Relativistic Hyperelastic continuous media in the mindset of Gauge Theory of High Energy Physics and Quantum Mechanics [33, 10]. The cornerstone of the theory is the representation of perfect matter by a vector-valued function $\Psi$, defined on the Universe and with values in a 3-dimensional vector space $V \simeq \mathbb{R}^3$. Then, the 3-dimensional body $\mathcal{B}$ of Classical Continuum Mechanics [68] is recovered as a compact submanifold with boundary of $V$ endowed with a mass measure $\mu$ that labels the material particles. Its preimage by $\Psi$, the Matter World tube, corresponds to the whole evolution of the considered matter in the Universe. Note that this description of continuous media is inverse to the one used in Classical Continuous Mechanics where the configurations are described by embeddings $p: \mathcal{B} \to \mathcal{E}$, where $\mathcal{E}$ is the three-dimensional space of Classical Mechanics. However, this description will be recovered once an observer (or time function) is introduced, which leads to the definition of a foliation of the Universe by spatial hypersurfaces (corresponding to the usual three-dimensional space $\mathcal{E}$ [31, 39]). Souriau's formulation of Hyperelasticity does not require a priori the definition of a time function. He can therefore express the consequences of demanding General Covariance for the matter Lagrangian. Assuming that the contribution of the matter Lagrangian density depends only on the zero order jet of the metric $g$ and on the first order jet $(\Psi, T\Psi)$ of the perfect matter field $\Psi$, Souriau's 1958 theorem [59] states that General Covariance implies that
this density depends only on the punctual value of $\Psi$ and of the secondary variable

$$ \mathbf{K} := T\Psi g^{-1} T\Psi^*, $$

named by him, the conformation, where $(\cdot)^*$ means the dual transpose.

**Theorem 1.1** (Souriau, 1958). Suppose that the Lagrangian defined on the 4-dimensional Universe,

$$ \mathcal{L}^{\text{matter}}[g, \Psi] = \int L_0 (g_{\mu\nu}, \Psi^I, \partial_{\lambda}\Psi^I) \text{vol}_g, $$

is general covariant, where $g$ is a Lorentzian metric, $\text{vol}_g$, the corresponding volume form, and $\Psi$ a matter field. Then, its Lagrangian density $L_0$ can be recast as $L_0 = L(\Psi, \mathbf{K})$, for some function $L$.

In the present work, we extend Souriau's theorem to a second order gradient theory in General Relativity. The starting point is the introduction of a second-order jet Lagrangian density, depending only on the two primary variables $g$ (the Universe metric) and $\Psi$ (the perfect matter field), and assumed to be general covariant [59, 61, 36, 6]. We define accordingly secondary variables (which are invariants of the diffeomorphism group) and playing the same role as Souriau's conformation, but for a relativistic second gradient theory. Among them are exhibited three novel gravitation/matter field coupling tensors (see subsection 3.2).

The central idea of this work is to find a systematic way to extract relativistic invariants from general covariant tensors, in the spirit of Souriau's definition of $\mathbf{K}$, which is a vector valued relativistic invariant. The main contribution of this work is the discovery that the correct formulation of the rest frame is a general covariant of the metric $g$ and the matter field $\Psi$. This allows first to produce a simple proof of Souriau's theorem and, then, to prove our result without sweating blood. There are no limit to extend our methods to higher order theories or to enrich the physical structure of the universe, but one needs a (perfect) matter field to make it work (this limits the scope of our technique to interior solutions, i.e. solutions inside matter, of General Relativity).

Finally, we show that the classical 3-dimensional second gradient theory can be derived from such a relativistic theory, as the Galilean approximation. Based on our General Covariance result, we address the issue of the status of the Principle of Objectivity in Classical (non-relativistic) Continuum Mechanics. Specifically, we show that the spatial gradient-type General Covariant secondary variables (modeling the deformation of solids) converge to objective quantities at the Galilean limit, while time derivative-type General Covariant secondary variables may converge to non-objective quantities at the Galilean limit.

The outline of the paper is as follows. In section 2, we recall some basic concepts in Variational Relativity: General Covariance, Lagrangian action functionals, matter field and matter rest frame. In section 3, we state and prove our main result, Theorem 3.3. We also provide an application to gradient fluids and a new and easy proof of Souriau's theorem. In section 4, we introduce the concept of time function, of observer frame, and the associated foliation by spacelike hypersurfaces (or so-called space-time structure). Our relativistic invariants are rewritten using this space-time decomposition and useful specifications of them are provided in the case of the Minkowski spacetime. Finally in section 5, we calculate their classical limits and discuss the question of objectivity (frame-indifference) of these limits. Besides, in order to be as self-contained as possible, we have added two appendices to summarize the main mathematical concepts and formulas used in this paper,

## 2. MATTER FIELDS IN VARIATIONAL RELATIVITY

The Universe is assumed to be a 4-dimensional orientable manifold $\mathcal{M}$, endowed with an hyperbolic metric $g$, of signature $(-+++)$.
2.1. **General Covariance.** The main postulate of General Relativity is precisely that *Physical laws must be independent of the choice of coordinates* [18]. This principle is known as *General Covariance*, or invariance by reparameterization. In order to lift the ambiguity created by the imprecise wording of this principle [54, 69], let us insist on what is meant here.

The Universe $\mathcal{M}$ is locally diffeomorphic to $\mathbb{R}^4$, via either coordinate charts

$$x: \mathcal{M} \to \mathbb{R}^4$$

or via parameterizations

$$\xi: \mathbb{R}^4 \to \mathcal{M},$$

also called *plots* by Souriau and Iglesias-Zemmour in a more general setting, see [62, 34]. Of course, parameterizations are the inverse of coordinates charts and *vice versa*. A reparameterization at the source (or *change of coordinates*) is a local diffeomorphism $\Phi$ from $\mathbb{R}^4$ to $\mathbb{R}^4$, whereas a reparameterization at the goal (or plainly a reparameterization) is a local diffeomorphism $\varphi$ from $\mathcal{M}$ to $\mathcal{M}$. Both changes of coordinates and reparameterizations act on either coordinates charts or parameterizations, by composition. However, whereas adopting the point of view of coordinates charts or of parameterizations is a strictly equivalent choice, it is not anymore the case for changes of coordinates and reparameterizations (see the figures and the explanations below).

<figure><img src="image_3.png" /><figcaption>FIGURE 1. A change of coordinates (left), and a reparameterization at the source (right).</figcaption></figure>

<figure><img src="image_4.png" /><figcaption>FIGURE 2. A reparameterization (at the goal).</figcaption></figure>

Firstly, there is no *canonical* equivalence between changes of coordinates and reparameterizations: one needs to fix an arbitrary (local) diffeomorphism $\mathbb{R}^4 \to \mathcal{M}$. Secondly and more importantly, even after identification, the actions of the two diffeomorphism groups are different. Changes of coordinates are involved in the process of gluing *local objects* into *global objects* such as tensor fields, whereas change of parameterizations are used to compare *global objects* at different positions by sending them in the same *local chart*.

One often aims at formulating theories using intrinsic formulas involving tensor fields (which is checked using changes of coordinates). General covariant tensor fields depend themselves on other tensor fields in a general covariant manner (which is checked using change of parameterizations).
Now that the principle is clearly stated, let us be a little more concrete. Variational Relativity, as formulated by Souriau [59, 61], involves primary variables (for instance, the metric $g$), and secondary variables, which depend on the primary variables in a covariant manner, for instance the Riemannian covariant derivative $\nabla^g$, the Riemann curvature tensor $R^g$, the Ricci tensor $\text{Ric}^g$ or the scalar curvature $R^g$ (see Appendix A). Acting in a covariant manner means that an action of the diffeomorphism group (denoted $\star$, see subsection A.2) on the primary variable $\mathbf{P}$ and the secondary variable $\mathbf{S}$ has been defined in such a way that

$$ \mathbf{S}(\varphi \star \mathbf{P}) = \varphi \star \mathbf{S}(\mathbf{P}), $$

for all local diffeomorphisms $\varphi$ of the Universe $\mathcal{M}$. We then say that $\mathbf{S}$ is a covariant of $\mathbf{P}$. A special case, which is of great importance, is when $\mathbf{S}$ is a scalar field or vector valued function and where the action of the diffeomorphism group on $\mathbf{S}$ is given by

$$ (\varphi \star \mathbf{S})(m) = \mathbf{S}(\varphi(m)). $$

In that case, we get

$$ (2.1) \qquad \mathbf{S}(\varphi \star \mathbf{P})(m) = \mathbf{S}(\mathbf{P})(\varphi(m)). $$

We shall then say that $\mathbf{S}$ is a relativistic invariant. Note that the scalar curvature $R^g$ is a relativistic invariant, but the Riemann curvature tensor $R^g$ and the Ricci tensor $\text{Ric}^g$ are not.

## 2.2. Variational Relativity.
We now apply the principle of General Covariance to the context of Lagrangian field theory of General Relativity [59, 61, 33]. In Lagrangian field theory, the primary variables are field variables defined on the Universe $\mathcal{M}$. The Lagrangian functional (or action)

$$ \mathcal{L}[\mathbf{P}] := \int L(j_m^n \mathbf{P}) \, \text{vol}_g $$

is defined using a Lagrangian density $L$ which depends on the punctual values (at $m \in \mathcal{M}$) of the primary variables $\mathbf{P}$ and their derivatives up to a given order $n$ (what is traditionally called their $n$th order jet and denoted by $j_m^n \mathbf{P}$). Usually, when the Lagrangian satisfies the principle of General Covariance, its Lagrangian density $L$ does not depend on the whole jet of $\mathbf{P}$ but rather on the $n$th order jet of $\mathbf{P}$ through secondary variables $\mathbf{S}$ which are covariants of the diffeomorphism group. For instance, in the initial variational formulation of General Relativity (in the vacuum), there is a unique primary variable, the metric $g$, and the Lagrangian density is given by

$$ L(j_m^2 g) = L(g_{\mu\nu}(m), \partial_{\kappa} g_{\mu\nu}(m), \partial_{\lambda\kappa}^2 g_{\mu\nu}(m)) = \frac{1}{2\kappa} (R^g(m) - \Lambda), $$

where $\kappa$ is the Einstein constant and $\Lambda$ is the cosmological constant. Here, $R^g$ is the scalar curvature of $g$ which is a secondary variable and a covariant of $g$, meaning that $R^{\varphi^*g} = \varphi^* R^g$ (see subsection A.2).

In this work, we will consider Lagrangian functionals $\mathcal{L}$ that depend on two primary variables: the Lorentzian metric $g$ of the Universe $\mathcal{M}$ and a vector valued function $\Psi: \mathcal{M} \to \mathbb{R}^3$, called the matter field [59, 61, 39]. In order to avoid unnecessary analytical difficulties and since, in practice, we do not require that Lagrangian densities $L$ are integrable over the whole manifold $\mathcal{M}$, usually not compact. Instead, Lagrangian densities are integrated only over relatively compact domains $\mathcal{U} \subset \mathcal{M}$ (and furthermore contained in a local chart). Therefore, we shall write

$$ \mathcal{L}_{\mathcal{U}}[g, \Psi] = \int_{\mathcal{U}} L(j_m^\ell g, j_m^n \Psi) \, \text{vol}_g, $$

to emphasize the dependence on $\mathcal{U}$.

Let us describe the principle of General Covariance for the Lagrangian $\mathcal{L}$ in more precise terms and formulate its consequences. Let $\varphi: \mathcal{U} \to \tilde{\mathcal{U}}$ be a diffeomorphism between two open sets $\mathcal{U}$ and $\tilde{\mathcal{U}}$ of the Universe $\mathcal{M}$ and let $T\varphi: T\mathcal{U} \to T\tilde{\mathcal{U}}$ be its tangent map. Then, the Lagrangian $\mathcal{L}$ is invariant by $\varphi$ if

$$ \mathcal{L}_{\mathcal{U}}[\varphi^*g, \varphi^*\Psi] = \mathcal{L}_{\tilde{\mathcal{U}}}[g, \Psi], $$
for every Lorentzian metrics $g$ on $\mathcal{M}$, every vector valued functions $\Psi: \mathcal{M} \to \mathbb{R}^3$, and every local diffeomorphism $\varphi: \mathcal{U} \to \tilde{\mathcal{U}}$. Here, $\varphi^*$ means the action by pullback of the local diffeomorphism $\varphi$ (see subsection A.2). One has

$$ \varphi^*g := (T\varphi)^*(g \circ \varphi)(T\varphi), \quad \varphi^*\Psi := \Psi \circ \varphi, $$

where $(\cdot)^*$ is the dual transpose. Using a change of variables, we hence observe that the Lagrangian is invariant by $\varphi$ if

$$ \int_{\mathcal{U}} L(j_m^\ell(\varphi^*g), j_m^n(\varphi^*\Psi)) \mathrm{vol}_{\varphi^*g} = \int_{\tilde{\mathcal{U}}} L(j_{\tilde{m}}^\ell g, j_{\tilde{m}}^n \Psi) \mathrm{vol}_g = \int_{\mathcal{U}} L(j_{\varphi(m)}^\ell g, j_{\varphi(m)}^n \Psi) \varphi^* \mathrm{vol}_g. $$

Since the open set $\mathcal{U}$ is arbitrary and $\mathrm{vol}_{\varphi^*g} = \varphi^*\mathrm{vol}_g$, one finally obtains the following condition on the Lagrangian density, for arbitrary local diffeomorphisms $\varphi$:

$$ (2.2) \qquad L(j_m^\ell(\varphi^*g), j_m^n(\varphi^*\Psi)) = L(j_{\varphi(m)}^\ell g, j_{\varphi(m)}^n \Psi). $$

General covariance of the Lagrangian functional being equivalent to condition (2.2), we get the following elementary result, on which our general reasoning in section 3 will rely.

**Lemma 2.1.** A Lagrangian functional is general covariant if and only if its density is a relativistic invariant. As a consequence, a Lagrangian which density depends only on relativistic invariants (definition (2.1)) is automatically general covariant.

**2.3. Matter field.** In Variational Relativity, following Souriau [59, 61], matter is described by the set of all lifelines (or World lines) of material points constitutive of the continuous medium under study. More precisely, one considers a smooth vector valued function

$$ \Psi: \mathcal{M} \to V \simeq \mathbb{R}^3. $$

The labels of these particles constitute a subset $\mathcal{B} \subseteq V$, in general assumed to be a 3-dimensional compact orientable manifold with boundary and called the body. It is further assumed that $\Psi$ is a submersion on

$$ \mathcal{W} := \Psi^{-1}(\mathcal{B}), $$

meaning that the linear tangent map

$$ T\Psi: T\mathcal{W} \to TV $$

is of rank 3 at each point of $\mathcal{W}$. Thus, $\mathcal{W}$ is fibered by the particles lifelines $\Psi^{-1}(\mathbf{X})$, for $\mathbf{X} \in \mathcal{B}$, and is called for this reason the body's World tube. It is moreover assumed (perfect matter assumption) that each 1-dimensional subspace ker $T_m\Psi$ of $T_m\mathcal{M}$ is time-like, meaning that the restriction of the metric $g$ to these subspaces is negative definite.

The body $\mathcal{B}$ is equipped, like in Classical Continuum Mechanics, with a mass measure $\mu$ (in our case, a volume form). Using the canonical coordinate system $(X^1)$ and the canonical metric $\mathbf{q} = (\delta_{IJ})$ of $V \simeq \mathbb{R}^3$, $\mu$ can be recast as

$$ \mu = \rho_o \mathrm{vol}_{\mathbf{q}}, \quad \mathrm{vol}_{\mathbf{q}} = \mathrm{d}X^1 \wedge \mathrm{d}X^2 \wedge \mathrm{d}X^3, $$

where $\rho_o$ is a positive function supported on the body and interpreted as a mass density on $\mathcal{B}$. The integral

$$ m = \int_{\mathcal{B}} \mu $$

is itself interpreted as the total mass of the continuous media under study.

The pullback $\Psi^*\mu$ by $\Psi$ of this mass measure $\mu$ is a 3-form defined on the 4-dimensional World tube $\mathcal{W} = \Psi^{-1}(\mathcal{B})$. Since $\Psi$ is assumed to be a submersion on $\mathcal{W}$, there exists a nowhere vanishing vector field $\mathbf{P}$ on $\mathcal{W}$, such that (see [59, 61, 39], cf. also (B.4))

$$ (2.3) \qquad i_{\mathbf{P}} \mathrm{vol}_g = \Psi^*\mu, $$

where $i_{\mathbf{P}}$ is the interior product (or left contraction) with $\mathbf{P}$. This vector field $\mathbf{P}$ is called the current of matter. It spans ker$(T\Psi)$ and is thus time-like. The rest mass density on the World tube $\mathcal{W}$ is the scalar

$$ \rho_r := \sqrt{-g(\mathbf{P}, \mathbf{P})}, $$
and we define the *quadrivelocity* $\mathbf{U}$ (this will be clarified with the introduction of a spacetime in section 5) as the unit time-like vector field on $\mathcal{W}$ such as

$$ (2.4) \qquad \mathbf{P} = \rho_r \mathbf{U}, \quad g(\mathbf{U}, \mathbf{U}) = -1. $$

*Remark 2.2* (Mass conservation, first expression). Since $d\mu = 0$ and $(\mathrm{div}^g \mathbf{P}) \mathrm{vol}_g = d(\mathbf{P} \mathrm{vol}_g) = d\Psi^*\mu = \Psi^* d\mu$, a first (kinetic) relativistic expression of mass conservation is

$$ \mathrm{div}^g \mathbf{P} = \mathrm{div}^g(\rho_r \mathbf{U}) = 0. $$

The *conformation tensor* $\mathbf{K}$ is a vector valued function defined on $\mathcal{M}$ with values in $S^2\mathbb{R}^3$ and given by

$$ (2.5) \qquad \mathbf{K} := T\Psi g^{-1}T\Psi^*, $$

where $T\Psi^*$ is the dual map (or transpose) of $T\Psi$. Note that $\mathbf{K}(m)$ is a symmetric contravariant tensor on $\mathbb{R}^3$ which is positive definite for each $m \in \mathcal{W}$ for perfect matter (see [41, Theorem p. 9] and [61, 39]).

We denote by $\Psi^I$ the components of the vector-valued function matter field $\Psi: \mathcal{M} \to \mathbb{R}^3$ in the canonical coordinate system $(X^I)$ of $\mathbb{R}^3$. In this coordinate system, the components of the conformation are

$$ K^{IJ} = g^{-1}(d\Psi^I, d\Psi^J). $$

Introducing the quadrivector field $\mathbf{W}$, defined implicitly on $\mathcal{M}$ by

$$ (2.6) \qquad i_{\mathbf{W}} \mathrm{vol}_g = \Psi^* \mathrm{vol}_q, $$

so that the current of matter can be written as $\mathbf{P} = (\rho_\circ \circ \Psi)\mathbf{W}$ on $\mathcal{W}$, and applying (B.5), we retrieve that

$$ g(\mathbf{P}, \mathbf{P}) = -(\rho_\circ \circ \Psi)^2 \det (g^{-1}(d\Psi^I, d\Psi^J)). $$

and the following expression of the rest mass density due to Souriau [61]

$$ (2.7) \qquad \rho_r = (\rho_\circ \circ \Psi) \sqrt{\det(\mathbf{K}\mathbf{q})}, $$

where $\det(\mathbf{K}\mathbf{q}) = \det(K^{IJ})$ is the determinant of the matrix $(K^{IJ})$.

*Remark 2.3* (Mass conservation, second expression). The expression (2.7) of the rest mass density is a second (integrated) relativistic expression of mass conservation. It is a generalization of the Classical Continuum Mechanics expression $\rho_\circ \circ \phi = \rho_0 / \sqrt{\det(C^{IJ})}$, where $\rho_0$ and $\rho$ are respectively, the mass densities on the 3D reference configuration $\Omega_0$ and the deformed configuration $\Omega = \phi(\Omega_0)$, and where $\mathbf{C}$ is the right Cauchy–Green tensor (defined on $\Omega_0$).

2.4. **Matter rest frame.** The vector field $\mathbf{W}$ defined implicitly on the Universe $\mathcal{M}$ by

$$ i_{\mathbf{W}} \mathrm{vol}_g = d\Psi^1 \wedge d\Psi^2 \wedge d\Psi^3 $$

is non-zero at each point $m \in \mathcal{M}$ at which $\Psi$ is a submersion. It is therefore non-vanishing on the world-tube $\mathcal{W}$.

In Appendix B.2, we prove that the covector field $\mathbf{W}^\flat = g\mathbf{W}$ is given by

$$ (2.8) \qquad \mathbf{W}^\flat := *_g(d\Psi^1 \wedge d\Psi^2 \wedge d\Psi^3), $$

where $*_g$ denotes the Hodge star operator for metric $g$, and therefore that $\mathbf{W}$ is orthogonal to the subspace spanned by the vector fields

$$ \mathrm{grad}^g \Psi^I = (d\Psi^I)^\sharp = g^{-1} d\Psi^I, \quad I = 1, 2, 3. $$

Since $\mathbf{W}$ is timelike on the world-tube $\mathcal{W}$ (perfect matter hypothesis), we conclude furthermore that $\langle \mathbf{W} \rangle^\perp$ is spacelike (see [41, Theorem p. 9]) and thus that the family

$$ (\mathbf{W}, \mathrm{grad}^g \Psi^1, \mathrm{grad}^g \Psi^2, \mathrm{grad}^g \Psi^3) $$

is a frame of $T_m\mathcal{M}$ at each point $m \in \mathcal{W}$.
In Appendix B.2, we show moreover that $W^b$ (and $W$) depends on the punctual value of the Lorentzian metric $g$ and on the first order jet of the matter field $\Psi$. It is furthermore general covariant, meaning that

$$W^b(\varphi^*g, \varphi^*\Psi, \varphi^*T\Psi) = \varphi^*W^b(g, \Psi, T\Psi),$$

for every local diffeomorphism $\varphi$ of $\mathcal{M}$.

**Remark 2.4.** Whereas the covector fields $d\Psi^I$ are naturally covariant under the group of general diffeomorphisms, the supplementary covector field $W^b = W^b(g, \Psi, T\Psi)$ changes sign when the orientation is reversed (see Appendix B.1 for more details).

This covariance property is inherited by the unit timelike vector field

$$U = \frac{W}{\sqrt{-g(W, W)}}$$

and by its covariant version $U^b$, since $g(W, W)$ is a relativistic invariant. Therefore, we shall introduce the coframe

$$\mathcal{F}_{\text{rest}}^{*} := (-U^{b}, E^{1}, E^{2}, E^{3}), \quad E^{I} := d\Psi^{I},$$

defined at each point $m \in \mathcal{W}$, and its dual frame $\mathcal{F}_{\text{rest}}$

$$\mathcal{F}_{\text{rest}} = (U, E_1, E_2, E_3).$$

One has thus for $I, J = 1, 2, 3$

$$(2.9) \qquad U^b(E_I) = 0 \quad \text{and} \quad d\Psi^I(E_J) = \delta^I_J.$$

This frame $\mathcal{F}_{\text{rest}}$ shall be referred to as the *matter rest frame*, or simply as the *rest frame*.

**Remark 2.5.** Since $T_m\Psi: T_m\mathcal{U} \to V$ is surjective, it has right inverses, which means linear mappings $L_m: V \to T_m\mathcal{U}$, such that $T_m\Psi \cdot L_m = \text{id}_V$. Each of them depends on the choice of a complement $S$ to $\ker T_m\Psi$ and will be denoted by $(T_m\Psi)_S^{-1}$. Among them, there is one which is “nicer” than the others, the one corresponding precisely to $(\ker T_m\Psi)^\perp$ (which is a complement of $\ker T_m\Psi$ because we have assumed that $\ker T_m\Psi$ is timelike on $\mathcal{U}$). This one will be called the *orthogonal right inverse* of $T_m\Psi$ and denoted by $(T_m\Psi)_{U^\perp}^{-1}$. Using this definition, one can observe that the vector fields $E_J$ can be written as

$$E_J = (T_m \Psi)_{U^\perp}^{-1} \cdot \frac{\partial}{\partial X^J},$$

introducing the canonical basis $(\partial/\partial X^I)$ on $V \simeq \mathbb{R}^3$. Indeed, we have

$$U^b(E_J) = 0, \quad \text{and} \quad E^I(E_J) = dX^I \cdot T_m\Psi \cdot (T_m\Psi)_{U^\perp}^{-1} \cdot \frac{\partial}{\partial X^J} = \delta^I_J.$$

We shall summarize all these considerations in the following theorem.

**Theorem 2.6.** Assume that $\Psi: \mathcal{M} \to V$ is a submersion on $\mathcal{W}$ and that $\ker T_m\Psi$ is timelike at each point $m \in \mathcal{W}$. Then, the frame

$$\mathcal{F}_{\text{rest}} = (U, E_1, E_2, E_3),$$

defined on $\mathcal{W}$, and its dual frame

$$\mathcal{F}_{\text{rest}}^{*} = (-U^{b}, E^{1}, E^{2}, E^{3}),$$

where $E^I := d\Psi^I$, depend only on the punctual value of the Lorentzian metric $g$ and of the first order jet of the matter field $\Psi$. Moreover, given an orientation-preserving local diffeomorphism $\varphi$ of $\mathcal{W}$, one has

$$\varphi^* \mathcal{F}_{\text{rest}}(g, \Psi, T\Psi) = \mathcal{F}_{\text{rest}}(\varphi^*g, \varphi^*\Psi, \varphi^*T\Psi),$$

$$\varphi^* \mathcal{F}_{\text{rest}}^*(g, \Psi, T\Psi) = \mathcal{F}_{\text{rest}}^*(\varphi^*g, \varphi^*\Psi, \varphi^*T\Psi).$$

In other words, the frame $\mathcal{F}_{\text{rest}}$ and its coframe $\mathcal{F}_{\text{rest}}^*$ are general covariant.
### 3. GENERAL COVARIANT LAGRANGIANS INVOLVING MATTER FIELDS

The introduction of the rest frame is essential to build relativistic invariants, similar to the components of the conformation $K$, in order to formulate a general covariant theory of gradient media in General Relativity.

**3.1. First Gradient Relativistic Hyperelasticity.** Relativistic Hyperelasticity has been formulated by Souriau in [59, 61] starting with a matter Lagrangian density

$$L_0 = L_0(g_{\mu\nu}(m), \Psi^I(m), \partial_{\lambda}\Psi^I(m)),$$

defined at each point $m \in \mathcal{M}$. It is a function of the punctual values $g(m)$ of the metric, of the matter field $\Psi(m)$ and of its first derivative $T_m\Psi$. Souriau has shown that General Covariance of the matter Lagrangian implies that its density can be recast as a function of $\Psi(m)$ and of the conformation $K(m)$ (Theorem 1.1). In order to illustrate the ideas of our new method—that can be extended straightforwardly to higher gradients— we first produce an alternative simpler proof of Souriau’s 1958 Theorem 1.1, but which is valid only on the World tube $\mathcal{W}$. This is common in General Relativity when one looks for interior solutions of Einstein’s equation (see [48, 49]), that is solutions defined inside matter.

*Proof of Theorem 1.1.* Starting with the variables $(g, \Psi, T\Psi)$, we first construct the coframe $\mathcal{F}_{rest}^*(g, \Psi, T\Psi)$ as in subsection 2.3. Since it amounts to add the covector field $U^b(g, \Psi, T\Psi)$ to $T\Psi$, one can safely replace the variables $(g, \Psi, T\Psi)$ by $(g, \Psi, \mathcal{F}_{rest}^*)$.

One can now replace the metric tensor $g$ by the matrix of the components of the cometric $g^{-1}$ in the coframe $\mathcal{F}_{rest}^*$. Taking account of (2.5), this amounts to replace the punctual value of the metric tensor $g$ by the one of the conformation tensor $K$, whose components ($K^{IJ}$) in the rest frame are relativistic invariants.

Hence, without still having used the general covariance of the Lagrangian, we have been able to recast the function $L_0(g, \Psi, T\Psi)$ as a new function $L(\Psi, K, \mathcal{F}_{rest}^*)$, depending on the relativistic invariants $\Psi$ and $K$, and on the general covariant coframe $\mathcal{F}_{rest}^*$. The general covariance of the Lagrangian density hence becomes

$$L(\Psi, K, \mathcal{F}_{rest}^*) = L(\Psi, K, \varphi^* \mathcal{F}_{rest}^*),$$

for any orientation-preserving local diffeomorphism $\varphi$ of $\mathcal{M}$.

Since the action of the (orientation-preserving) diffeomorphism group on the punctual value of the rest frame $\mathcal{F}_{rest}^*$ involves only the standard action of the orientation-preserving linear group $\varphi \in GL^+(4, \mathbb{R})$ which acts transitively on the oriented bases of $\mathbb{R}^4$, we conclude that $L$ does not depend on $\mathcal{F}_{rest}^*$. This achieves the proof. $\square$

**Remark 3.1.** Note a posteriori that the result is still true if one considers covariance under the group of general diffeomorphisms instead of only orientation-preserving diffeomorphisms, because the vector field $U$ is not involved in the formation of the relativistic invariants $\Psi$ and $K^{IJ}$. This will not anymore be the case for higher order gradients.

**Remark 3.2.** The choice to select the components of the cometric rather than of the metric is natural. Indeed, since $T\Psi$ is not invertible, one can pushforward a contravariant tensor, but not a covariant tensor (see subsection A.2).

**3.2. Second Gradient Relativistic Hyperelasticity.** We now seek for a general covariant formulation of Relativistic Gradient Hyperelasticity and start with a matter Lagrangian density of general expression

$$L_0(g_{\mu\nu}, \partial_{\nu}g_{\lambda\mu}, \partial_{\mu}\partial_{\nu}g_{\kappa\lambda}, \Psi^I, \partial_{\lambda}\Psi^I, \partial_{\lambda}\partial_{\mu}\Psi^I),$$

defined at each point $m \in \mathcal{M}$ of the Universe. This Lagrangian density is a function of the second order jet of the metric $g$ and of the second order jet of the matter field $\Psi$. We shall assume, furthermore, that $\Psi$ is a submersion and $\ker(T\Psi)$ is timelike.
**Theorem 3.3.** Consider the Lagrangian

$$ \mathcal{L}^{\text{matter}}[g, \Psi] = \int L_0(g_{\mu\nu}, \partial_{\nu}g_{\lambda\mu}, \partial_{\mu}\partial_{\nu}g_{\kappa\lambda}, \Psi^I, \partial_{\mu}\Psi^I, \partial_{\mu}\partial_{\nu}\Psi^I) \text{vol}_g $$

depending on the field variables $g$, a Lorentzian metric, and $\Psi: \mathcal{M} \to V$, a submersion such that $\ker(T\Psi)$ is timelike. Then, if $\mathcal{L}^{\text{matter}}$ is general covariant under the group of orientation-preserving diffeomorphisms, its Lagrangian density $L_0$ can be recast, at each point of the Wold tube $\mathcal{W}$, as

$$ L_0 = L ([g^{-1}]_{\mathcal{F}_{\text{rest}}}, [\mathbf{R}^g]_{\mathcal{F}_{\text{rest}}}, \Psi, [\text{Hess}^g \Psi]_{\mathcal{F}_{\text{rest}}}), $$

for some function $L$, where $\mathbf{R}^g$ is the curvature tensor, $\text{Hess}^g \Psi$ is the Hessian tensor of the vector valued function $\Psi$, and where $[\mathbf{t}]_{\mathcal{F}_{\text{rest}}}$ denotes the set of components of a tensor field $\mathbf{t}$ in $\mathcal{F}_{\text{rest}}$.

**Remark 3.4.** Note that, in contrast with **Theorem 1.1**, the result does not extend to covariance under the full group of general diffeomorphisms. Components that involve the time-like vector field $\mathbf{U}$ an even number of time are invariant under the group of general diffeomorphisms, but components that involve it an odd number of times change sign when the orientation is reversed.

**Proof.** The proof is very similar to our new proof of **Theorem 1.1**.

(1) Starting with the variables

$$ (g_{\mu\nu}, \partial_{\nu}g_{\lambda\mu}, \partial_{\mu}\partial_{\nu}g_{\kappa\lambda}, \Psi^I, \partial_{\mu}\Psi^I, \partial_{\mu}\partial_{\nu}\Psi^I), $$

a first change of variables consists in substituting the first and second derivatives of $g_{\mu\nu}$ by the Christoffel symbols of the corresponding Levi-Civita connection and their derivatives (see **Appendix A**). The variables are now

$$ (g_{\mu\nu}, \Gamma^{\lambda}_{\mu\nu}, \partial_{\nu}\Gamma^{\kappa}_{\lambda\mu}, \Psi^I, \partial_{\mu}\Psi^I, \partial_{\mu}\partial_{\nu}\Psi^I). $$

(2) The Levi-Civita connection allows one to define the metric covariant derivative $\nabla^g$ of a tensor field. In particular, one can define the Hessian of the scalar function $\Psi^I$ (see **Appendix A**)

$$ \text{Hess}^g \Psi^I := \nabla^g d\Psi^I, $$

a second-order covariant symmetric tensor field, with components

$$ (\text{Hess}^g \Psi)^I_{\mu\nu} = \partial_{\mu}\partial_{\nu}\Psi^I - \Gamma^{\rho}_{\mu\nu}\partial_{\rho}\Psi^I. $$

This allows us to replace (using a change of variables) the second order derivatives of $\Psi^I$ by the components of its covariant Hessian. The variables are now

$$ (g_{\mu\nu}, \Gamma^{\lambda}_{\mu\nu}, \partial_{\nu}\Gamma^{\kappa}_{\lambda\mu}, \Psi^I, \partial_{\mu}\Psi^I, (\text{Hess}^g \Psi)^I_{\mu\nu}). $$

(3) So far, we have not used the general covariance hypothesis. At this stage, we have replaced most of the jet variables by the local components of general covariant tensors, at the exception of $\Gamma^{\lambda}_{\mu\nu}$ and $\partial_{\nu}\Gamma^{\kappa}_{\lambda\mu}$. We will now use the general covariance of the Lagrangian and the normalization result **Corollary A.8** established in **subsection A.3** in order to replace these remaining jet variables by components of general covariant tensors.

Consider the subgroup of $\text{Diff}^+(\mathcal{M})$ made of diffeomorphisms with trivial linear part, i.e.

$$ \{\varphi \in \text{Diff}(\mathcal{M}) : \varphi(m) = m \text{ and } T_m\varphi = \text{id}\}. $$

If $\mathbf{t}(g, \Psi)$ is a general covariant tensor field, then for such a diffeomorphism $\varphi$, one has the punctual invariance property

$$ \mathbf{t}(\varphi^*g, \varphi^*\Psi) = \varphi^*\mathbf{t}(g, \Psi) = \mathbf{t}(g, \Psi), $$

because punctually diffeomorphisms act on tensors via their linear parts. By **Corollary A.8**, there exists such a diffeomorphism with

$$ \Gamma^{\lambda}_{\mu\nu}(\varphi^*g) = 0, \quad \text{and} \quad \partial_{\nu}\Gamma^{\kappa}_{\lambda\mu}(\varphi^*g) = -\frac{1}{3}(R^{\kappa}_{\lambda\mu\nu}(\varphi^*g) + R^{\kappa}_{\mu\lambda\nu}(\varphi^*g)). $$
Since all other variables are fixed by $\varphi$, under the covariance assumption, one can replace the variables $\Gamma^\lambda_{\mu\nu}$ by 0 and $\Gamma^\kappa_{\lambda\mu,\nu} = \partial_\nu \Gamma^\kappa_{\lambda\mu}$ by the components $R^\kappa_{\lambda\mu\nu}$ of the curvature tensor $\mathbf{R}^g$ (which is itself a function of $\Gamma^\kappa_{\lambda\mu}$ and of $\partial_\nu \Gamma^\kappa_{\lambda\mu}$). The variables are now

$$ (g, \mathbf{R}^g, \Psi, T\Psi, \mathbf{Hess}^g \Psi). $$

(4) As in the proof of Theorem 1.1, we can replace (using the metric $g$) the variables $\partial_\mu \Psi^I$ by the coframe $\mathcal{F}_{\text{rest}}^*$, or equivalently, by the frame $\mathcal{F}_{\text{rest}}$. Finally, we can replace all the (general covariant) tensor fields by their components in the frame $\mathcal{F}_{\text{rest}}$, which are relativistic invariants. The variables are now

$$ ([g^{-1}]_{\mathcal{F}_{\text{rest}}}, [\mathbf{R}^g]_{\mathcal{F}_{\text{rest}}}, \Psi, \mathcal{F}_{\text{rest}}, [\mathbf{Hess}^g \Psi]_{\mathcal{F}_{\text{rest}}}). $$

(5) To sum up we have recast $L_0$ as a new Lagrangian density

$$ L_0 = L([g^{-1}]_{\mathcal{F}_{\text{rest}}}, [\mathbf{R}^g]_{\mathcal{F}_{\text{rest}}}, \Psi, \mathcal{F}_{\text{rest}}, [\mathbf{Hess}^g \Psi]_{\mathcal{F}_{\text{rest}}}), $$

for some function $L$, depending on some relativistic invariants, and on the general covariant frame $\mathcal{F}_{\text{rest}}^*$. We conclude, as in the proof of Theorem 1.1, that General Covariance implies that $L$ does not depend on $\mathcal{F}_{\text{rest}}$ and thus that

$$ L_0 = L([g^{-1}]_{\mathcal{F}_{\text{rest}}}, [\mathbf{R}^g]_{\mathcal{F}_{\text{rest}}}, \Psi, [\mathbf{Hess}^g \Psi]_{\mathcal{F}_{\text{rest}}}), $$

which is obviously a relativistic invariant. This concludes the proof. $\square$

We shall now detail these components, using the orthogonal 3+1 decomposition relative to the timelike vector field $\mathbf{U}$ (see [39, Appendix A]). Note first that

$$ (3.1) \qquad [g^{-1}]_{\mathcal{F}_{\text{rest}}} = \begin{bmatrix} -1 & 0 \\ 0 & \mathbf{K} \end{bmatrix}, $$

where $\mathbf{K}$ is the conformation tensor introduced by Souriau. To express the components of the Hessian, we introduce the speed of light $c$ and some normalizations (which will be justified once a space-time structure is introduced). We write thus

$$ (3.2) \qquad [\mathbf{Hess}^g \Psi^I]_{\mathcal{F}_{\text{rest}}} = \begin{bmatrix} -c^{-2}A^I & -c^{-1}\mathbf{L}^I \\ -c^{-1}(\mathbf{L}^I)^T & -\mathbf{H}^I \end{bmatrix}. $$

Here, the matrix-by-block representation (3.2) corresponds to the orthogonal decomposition of the Hessian tensor $\mathbf{Hess}^g \Psi^I$ relative to $\mathbf{U}$. In tensorial notation, it is written as

$$ \mathbf{Hess}^g \Psi = -\frac{1}{c^2} \mathbf{A} \otimes \mathbf{U}^b \otimes \mathbf{U}^b + \frac{2}{c} (\mathbf{L} T \Psi \otimes \mathbf{U}^b)^{(23)} - \mathbf{H}(\cdot, T\Psi, T\Psi), $$

where (23) means the symmetrization over the second and third factors. In components, we get

$$ (\mathbf{Hess}^g \Psi^I)_{\mu\nu} = -\frac{1}{c^2} A^I U_\mu U_\nu + \frac{1}{c} L^I_J (T \Psi^J_\mu U_\nu + T \Psi^J_\nu U_\mu) - H^I_{JK} T \Psi^J_\mu T \Psi^K_\nu, $$

where $U_\mu = (\mathbf{U}^b)_\mu$.

• The first order tensor $\mathbf{A} = (A^I)$ is homogeneous to an acceleration and will be called the material relativistic acceleration. It has 3 independent components,

$$ A^I := -c^2 (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{U}) = -c^2 (\nabla^g_{\mathbf{U}} d \Psi^I) \cdot \mathbf{U}, $$

and recast (since $d \Psi^I(\mathbf{U}) = 0$) as

$$ A^I = d \Psi^I \cdot (\nabla^g_{c\mathbf{U}} c \mathbf{U}), $$

or, in tensorial notation, as

$$ (3.3) \qquad \mathbf{A} = T\Psi \cdot \nabla^g_{c\mathbf{U}} (c\mathbf{U}). $$
• The second order tensor $\mathbf{L} = (L^I_J)$ is homogeneous to a spatial gradient of velocity and will be called the material relativistic velocity gradient. It has 9 independent components,

$$ (3.4) \qquad L^I_J = (\mathbf{L}^I)_J := -c (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{E}_J) = -c \nabla^g_{\mathbf{E}_J} (d\Psi^I) \cdot \mathbf{U}, $$

and recast (since $d\Psi^I(\mathbf{U}) = 0$) as

$$ L^I_J = d\Psi^I \cdot \nabla^g_{\mathbf{E}_J} (c\mathbf{U}), $$

or, in tensorial notation, as

$$ (3.5) \qquad \mathbf{L} = T\Psi (\nabla^g c\mathbf{U}) (T\Psi)^{-1}_{\mathbf{U}^{\perp}}. $$

• The third order tensor $\mathbf{H} = (H^I_{JK})$ is homogeneous to a strain gradient and will be called the material relativistic strain gradient. It satisfies the index symmetry $H^I_{JK} = H^I_{KJ}$ and has hence 18 independent components,

$$ (3.6) \qquad H^I_{JK} = (\mathbf{H}^I)_{JK} := -(\mathbf{Hess}^g \Psi^I)(\mathbf{E}_J, \mathbf{E}_K) = -\nabla^g_{\mathbf{E}_K} (d\Psi^I) \cdot \mathbf{E}_J, $$

and recast (since $d\Psi^I(\mathbf{E}_J) = \delta^I_J$) as

$$ H^I_{JK} = d\Psi^I \cdot \nabla^g_{\mathbf{E}_K} (\mathbf{E}_J) = d\Psi^I \cdot \nabla^g_{\mathbf{E}_J} (\mathbf{E}_K), $$

or, in tensorial notation, as

$$ (3.7) \qquad \mathbf{H} = -(\nabla^g T\Psi) (\cdot, (T\Psi)^{-1}_{\mathbf{U}^{\perp}}, (T\Psi)^{-1}_{\mathbf{U}^{\perp}} \cdot) = T\Psi \cdot \nabla^g_{\mathbf{E}_K} (\mathbf{E}_J) \otimes dX^J \otimes dX^K. $$

We will not use the components of the curvature tensor $\mathbf{R}^g$ for applications in this work. However, the data contained in $[\mathbf{R}^g]_{\mathcal{F}_{\text{rest}}}$ will be reorganized into three new relativistic gravitation/matter coupling tensors. These tensors $\mathbf{M}_p$ represent new coupling invariants for gravitation/matter and could be useful at the astrophysics scale, for the study of neutron stars, of black holes and of dark energy/matter. They vanish for a flat spacetime such as the Minkowski spacetime of Special Relativity. These relativistic invariants are defined as follows.

• The fourth order tensor $\mathbf{M}_4 = (M^I_{JKL})$ is given, in components, by

$$ M^I_{JKL} := \mathbf{R}^g(d\Psi^I, \mathbf{E}_J, \mathbf{E}_K, \mathbf{E}_L) = R^\kappa_{\lambda\mu\nu} T\Psi^I_\kappa E^\lambda_J E^\mu_K E^\nu_L, $$

or, in tensorial notation, by

$$ \mathbf{M}_4 = \mathbf{R}^g(T\Psi^* \cdot, (T\Psi)^{-1}_{\mathbf{U}^{\perp}}, (T\Psi)^{-1}_{\mathbf{U}^{\perp}} \cdot, (T\Psi)^{-1}_{\mathbf{U}^{\perp}} \cdot). $$

• The third order tensor $\mathbf{M}_3 = (M^I_{JK})$ is given, in components, by

$$ M^I_{JK} := c \mathbf{R}^g(d\Psi^I, \mathbf{E}_J, \mathbf{E}_K, \mathbf{U}) = c R^\kappa_{\lambda\mu\nu} T\Psi^I_\kappa E^\lambda_J E^\mu_K U^\nu, $$

or, in tensorial notation, by

$$ \mathbf{M}_3 = c \mathbf{R}^g(T\Psi^* \cdot, (T\Psi)^{-1}_{\mathbf{U}^{\perp}}, (T\Psi)^{-1}_{\mathbf{U}^{\perp}} \cdot, \mathbf{U}). $$

• The second order tensor $\mathbf{M}_2 = (M^I_J)$ is given, in components, by

$$ M^I_J := \mathbf{R}^g(d\Psi^I, c\mathbf{U}, \mathbf{E}_J, c\mathbf{U}) = c^2 R^\kappa_{\lambda\mu\nu} T\Psi^I_\kappa U^\lambda E^\mu_J U^\nu, $$

or, in tensorial notation, by

$$ \mathbf{M}_2 = c^2 \mathbf{R}^g(T\Psi^* \cdot, \mathbf{U}, (T\Psi)^{-1}_{\mathbf{U}^{\perp}} \cdot, \mathbf{U}). $$

**Remark 3.5.** The Riemann tensor $g\mathbf{R}^g = (R_{\kappa\lambda\mu\nu})$ having the following symmetries

$$ R_{\kappa\lambda\mu\nu} = R_{\mu\nu\kappa\lambda} = -R_{\lambda\kappa\mu\nu} = -R_{\kappa\lambda\nu\mu}, \quad \text{and} \quad R_{\kappa\lambda\mu\nu} + R_{\lambda\mu\kappa\nu} + R_{\mu\kappa\lambda\nu} = 0, $$

and 20 independent components in dimension 4, there are thus 6 independent components for $\mathbf{M}_4$, 8 for $\mathbf{M}_3$, and 6 for $\mathbf{M}_2$.

Reformulating **Theorem 3.3** using these new secondary variables, we get the following corollary.
**Corollary 3.6.** If the Lagrangian

$$ \mathcal{L}^{\text{matter}}(g, \Psi) = \int L_0(g_{\mu\nu}, \partial_{\nu}g_{\lambda\mu}, \partial_{\mu}\partial_{\nu}g_{\kappa\lambda}, \Psi^I, \partial_{\mu}\Psi^I, \partial_{\mu}\partial_{\nu}\Psi^I) \text{vol}_g $$

is general covariant under the group of orientation-preserving diffeomorphisms, then, its Lagrangian density $L_0$ can be written as a function

$$ L_0 = L(\Psi, \mathbf{K}, \mathbf{M}_2, \mathbf{M}_3, \mathbf{M}_4, \mathbf{A}, \mathbf{L}, \mathbf{H}) $$

of the vector valued relativistic invariants $\Psi, \mathbf{K}, \mathbf{M}_2, \mathbf{M}_3, \mathbf{M}_4, \mathbf{A}, \mathbf{L}, \mathbf{H}$.

**Remark 3.7.** In the spirit of Remark 3.2, one can formulate alternative relativistic invariants by using the contravariant forms $(\mathbf{Hess}^g \Psi)^{\sharp}$ and $(\mathbf{R}_g)^{\sharp}$ of the tensor fields $\mathbf{Hess}^g \Psi$ and $\mathbf{R}_g$. These alternative invariants are the components of $(\mathbf{Hess}^g \Psi)^{\sharp}$ and $(\mathbf{R}_g)^{\sharp}$—up to normalization factors— in the rest coframe $\mathcal{F}_{\text{rest}}^{\star} = (-\mathbf{U}^{\flat}, d\Psi^I)$. For instance, instead of the relativistic invariants $H_{JK}^I$ (defined by (3.6)), one may introduce

$$ H^{IJK} := -(\mathbf{Hess}^g \Psi^I)^{\sharp}(d\Psi^J, d\Psi^K) = -\mathbf{Hess}^g \Psi^I(\mathrm{grad}^g \Psi^J, \mathrm{grad}^g \Psi^K). $$

But since $\mathrm{grad}^g \Psi^J = K^{IP} \mathbf{E}_P$, we get

$$ H^{IJK} = K^{JP} K^{KQ} H_{PQ}^{I} = (\mathbf{H}^{\sharp}_{\mathbf{K}^{-1}})^{IJK}, $$

where the inverse of the conformation, $\mathbf{K}^{-1}$, is used as a metric to raise indices on the vector space $V$. The $H^{IJK}$ are the components of the contravariant material relativistic strain gradient,

$$ \mathbf{H}^{\sharp}_{\mathbf{K}^{-1}} = (\mathbf{Hess}^g \Psi)^{\sharp}(\cdot, T\Psi^*, T\Psi^*). $$

Similarly, instead of the relativistic invariants $L_J^I$ (defined by (3.4)) one has

$$ L^{IJ} := c(\mathbf{Hess}^g \Psi^I)^{\sharp}(-\mathbf{U}^{\flat}, d\Psi^J) = -c \mathbf{Hess}^g \Psi^I(\mathbf{U}, \mathrm{grad}^g \Psi^J) = L_P^I K^{PJ}, $$

which are the components of the contravariant material relativistic velocity gradient

$$ \mathbf{L}^{\sharp}_{\mathbf{K}^{-1}} = \mathbf{L} \cdot \mathbf{K}. $$

One also has

$$ M^{IJKL} := (\mathbf{R}^g)^{\sharp} (d\Psi^I, d\Psi^J, d\Psi^K, d\Psi^L) = K^{JP} K^{KQ} K^{LR} M^I_{PQR}, $$

$$ M^{IJK} := -c (\mathbf{R}^g)^{\sharp} (d\Psi^I, d\Psi^J, d\Psi^K, -\mathbf{U}^{\flat}) = K^{JP} K^{KQ} M^I_{PQ}, $$

$$ M^{IJ} := c^2 (\mathbf{R}^g)^{\sharp} (d\Psi^I, \mathbf{U}^{\flat}, d\Psi^K, \mathbf{U}^{\flat}) = K^{JP} M^I_P. $$

which are the components of $(\mathbf{M}_4)^{\sharp}_{\mathbf{K}^{-1}}$, $(\mathbf{M}_3)^{\sharp}_{\mathbf{K}^{-1}}$ and $(\mathbf{M}_2)^{\sharp}_{\mathbf{K}^{-1}} = \mathbf{M}_2 \cdot \mathbf{K}$, respectively.

**3.3. Gradient Fluids.** The classical (3D) gradient theory for fluids [27, 30] involves only the secondary variable $\rho$ (the fluid mass density) and considers its gradient $\mathrm{grad} \rho$, or $\mathrm{grad} \ln \rho$, as the main variable for constitutive modeling. In the present relativistic framework, we shall consider, as the fundamental secondary variable, the rest mass density

$$ (3.8) \qquad \rho_r(g, \Psi, T\Psi) = (\rho_\circ \circ \Psi) \sqrt{\det(K^{IJ})}, \quad \text{where} \quad \mathbf{K} = T\Psi g^{-1} T\Psi^*, $$

and $\rho_\circ$ is the mass density on the body $\mathcal{B}$ (a parameter). The function $\rho_r$ is only defined on the World tube $\mathcal{W}$ and can be used only to describe interior solutions. The vector field $\mathrm{grad}^g(\ln \rho_r)$ is a function of the metric $g$ and the second order jet of the matter field $\Psi$. Note that $\rho_r$ is a relativistic invariant but $\mathrm{grad}^g(\ln \rho_r)$ is not; it is only a relativistic covariant, meaning that

$$ \mathrm{grad}^g(\ln \rho_r)(\varphi^* g, \varphi^* \Psi, \varphi^* T\Psi) = \varphi^* (\mathrm{grad}^g(\ln \rho_r)(g, \Psi, T\Psi)), $$

for every (local) diffeomorphism $\varphi$. However, the components of this vector field in the matter frame $\mathcal{F}_{\text{rest}}$

$$ h^0 := -\mathbf{U}^{\flat} \cdot \mathrm{grad}^g(\ln \rho_r) = -d(\ln \rho_r) \cdot \mathbf{U}, \qquad h^I := d\Psi^I \cdot \mathrm{grad}^g(\ln \rho_r) = d(\ln \rho_r) \cdot \mathrm{grad}^g \Psi^I $$
are relativistic invariants. This means that one can formulate a general covariant gradient theory for relativistic fluids using the following matter Lagrangian density

$$L = \rho_r c^2 + E(\rho_r, h^0, h^I),$$

where the pure mass term $\rho_r c^2$ models dust [59, 61].

By corollary 3.6, the Lagrangian density $L$ should be expressible using the relativistic invariants $\Psi, K, A, L, H$. This is indeed the case, and we will summarize this in the following lemma.

**Lemma 3.8.** Let $(X^I)$ be the canonical coordinate system on $V = \mathbb{R}^3$. Given a mass measure

$$\mu = \rho_\circ dX^1 \wedge dX^2 \wedge dX^3$$

on the body $\mathcal{B}$, we have

$$h^0 = -\frac{1}{c} \mathrm{tr} \mathbf{L} \quad \text{and} \quad \mathbf{h} = (\mathrm{d}(\ln \rho_\circ) \circ \Psi) \cdot \mathbf{K} - (\mathrm{tr}_{12} \mathbf{H}) \cdot \mathbf{K},$$

where $\mathbf{h} = (h^I)$, $\mathbf{K} = (K^{IJ})$ is the conformation and $\mathrm{tr}_{12} \mathbf{H} = (H^J_{JI})$.

**Remark 3.9.** For an homogeneous fluid ($\rho_\circ$ = constant), we get simply

$$\mathbf{h} = -(\mathrm{tr}_{12} \mathbf{H}) \cdot \mathbf{K}, \quad h^I = -H^J_{JI} K^{LI}.$$

**Proof of lemma 3.8.** By (3.8), we have

$$\mathrm{d}(\ln \rho_r) = \mathrm{d}(\ln \rho_\circ \circ \Psi) + \frac{1}{2} \mathrm{d}(\ln \det(K^{IJ})).$$

But

$$\mathrm{d}(\ln \det(K^{IJ})) = (K^{-1})_{IJ} \mathrm{d}(K^{IJ}).$$

Now, for any vector field $\mathbf{X}$, one has

$$\begin{aligned} \mathrm{d} K^{IJ} \cdot \mathbf{X} &= \nabla_{\mathbf{X}}^g (g^{-1}(\mathrm{d}\Psi^I, \mathrm{d}\Psi^J)) \\ &= g^{-1}(\nabla_{\mathbf{X}}^g \mathrm{d}\Psi^I, \mathrm{d}\Psi^J) + g^{-1}(\mathrm{d}\Psi^I, \nabla_{\mathbf{X}}^g \mathrm{d}\Psi^J) \\ &= \mathbf{Hess}^g \Psi^I(\mathbf{X}, \mathrm{grad}^g \Psi^J) + \mathbf{Hess}^g \Psi^J(\mathbf{X}, \mathrm{grad}^g \Psi^I). \end{aligned}$$

But

$$\mathbf{X} = X^0 \mathbf{U} + X^L \mathbf{E}_L, \quad \mathrm{grad}^g \Psi^I = \mathbf{K}^{IM} \mathbf{E}_M,$$

with

$$\mathbf{Hess}^g \Psi^I(\mathbf{U}, \mathbf{E}_M) = -\frac{1}{c} L^I_M, \quad \mathbf{Hess}^g \Psi^I(\mathbf{E}_L, \mathbf{E}_M) = -H^I_{LM},$$

and thus

$$\mathbf{Hess}^g \Psi^I(\mathbf{X}, \mathrm{grad}^g \Psi^J) = -\frac{1}{c} X^0 K^{JM} L^I_M - X^L K^{JM} H^I_{LM}.$$

We get therefore

$$\begin{aligned} \mathrm{d}(\ln \det(K^{IJ})) \cdot \mathbf{X} &= (K^{-1})_{IJ} \mathrm{d}(K^{IJ}) \cdot \mathbf{X} \\ &= (K^{-1})_{IJ} \left( -\frac{1}{c} X^0 K^{JM} L^I_M - X^L K^{JM} H^I_{LM} \right) \\ &\quad + (K^{-1})_{IJ} \left( -\frac{1}{c} X^0 K^{IM} L^J_M - X^L K^{IM} H^J_{LM} \right) \\ &= -2 \left( \frac{1}{c} X^0 L^M_M + X^L H^M_{LM} \right), \end{aligned}$$

from which we deduce that

$$\frac{1}{2} \mathrm{d}(\ln \det(K^{IJ})). \mathbf{U} = -\frac{1}{c} L^M_M,$$

and

$$\frac{1}{2} \mathrm{d}(\ln \det(K^{IJ})). \mathrm{grad}^g \Psi^N = -K^{NL} H^M_{LM} = -H^M_{ML} K^{LN}.$$
By the way, we have

$$d(\ln \rho_{\circ} \circ \Psi).\mathbf{U} = 0, \quad \text{and} \quad d(\ln \rho_{\circ} \circ \Psi).\operatorname{grad}^g \Psi^N = (d \ln \rho_{\circ})_L K^{LN}.$$

We get thus finally

$$h^0 = -\frac{1}{c} \operatorname{tr} \mathbf{L} \quad \text{and} \quad \mathbf{h} = h^N \mathbf{E}_N = ((d \ln \rho_{\circ}) \circ \Psi) \cdot \mathbf{K} - (\operatorname{tr}_{12} \mathbf{H}) \cdot \mathbf{K}. \quad \square$$

*Remark 3.10.* Different relativistic invariants can be obtained, if one starts with $d(\ln \rho_r)$ rather than $\operatorname{grad}^g(\ln \rho_r)$. In that case, one gets

$$k_0 = \operatorname{tr} \mathbf{L}, \quad \text{and} \quad \mathbf{k} = (d \ln \rho_{\circ}) \circ \Psi - \operatorname{tr}_{12} \mathbf{H}.$$

## 4. TIME, SPACE AND OBSERVER FRAMES

4.1. **Observer.** In General Relativity, in order to observe the motion of matter, it is necessary to introduce an arbitrary time function $\hat{t}$ on the Universe $(\mathcal{M}, g)$, that is a smooth (local) submersion

$$\hat{t}: \mathcal{M} \to \mathbb{R},$$

defined on $\mathcal{M}$, whose gradient is everywhere time-like. This function defines a foliation (called a spacetime structure or (3 + 1)-structure [3, 70, 31]) of $\mathcal{M}$ by spacelike hypersurfaces

$$\mathcal{E}_t := \{m \in \mathcal{M}; \hat{t}(m) = t\}.$$

Let $\mathbf{N}$ be the future-oriented (i.e. $g(\mathbf{N}, \operatorname{grad}^g \hat{t}) < 0$) normal vector field to $\mathcal{E}_t$. Then, the orthogonal decomposition relative to $\mathbf{N}$ of the metric $g$ is given by

$$g = -\mathbf{N}^{\flat} \otimes \mathbf{N}^{\flat} + \mathbf{g}, \quad \text{with} \quad \mathbf{g}\mathbf{N} = 0.$$

The degenerate metric $\mathbf{g}$, when restricted to each hypersurface $\mathcal{E}_t$, is a Riemannian metric and coincides with the restriction of the Lorentzian metric $g$ to $\mathcal{E}_t$ [41, Page 6].

We shall assume furthermore that $g(\mathbf{U}, \operatorname{grad}^g \hat{t}) < 0$, where the timelike vector field $\mathbf{U}$ was defined in subsection 2.3 and is related to matter. This means that $\operatorname{grad}^g \hat{t}$ and $\mathbf{U}$ define the same time-orientation. This allows to define the generalized *Lorentz factor* as the scalar field [31]

$$\gamma := -g(\mathbf{U}, \mathbf{N}),$$

and the *spatial velocity* $\mathbf{u}$, defined by the orthogonal decomposition of $\mathbf{U}$ relative to $\mathbf{N}$ (see [59, 61, 31, 39]), given by

$$(4.1) \qquad \mathbf{U} = \gamma \left( \mathbf{N} + \frac{\mathbf{u}}{c} \right), \quad \text{with} \quad g(\mathbf{u}, \mathbf{N}) = 0.$$

Since $g(\mathbf{U}, \mathbf{U}) = g(\mathbf{N}, \mathbf{N}) = -1$, one recovers the familiar expression

$$\gamma = \frac{1}{\sqrt{1 - \frac{\|\mathbf{u}\|_g^2}{c^2}}}.$$

Let us now consider the spacelike hypersurfaces of the World tube

$$\Omega_t := \mathcal{W} \cap \mathcal{E}_t,$$

that play the same role as the configuration manifolds in Classical Continuum Mechanics. Note that whereas the restriction $\Psi_t: \Omega_t \to \mathcal{B}$ is not assumed (a priori) bijective, its differential $T\Psi_t$ is always a linear isomorphism. Since $\mathcal{E}_t$ is orthogonal to $\mathbf{N}$, at a point $m$ of $\Omega_t$ one has the orthogonal decomposition

$$T\Psi = -(T\Psi \cdot \mathbf{N}) \otimes \mathbf{N}^{\flat} + T\Psi_t,$$

where $T\Psi_t$ is the restriction of $T\Psi$ to the spacelike vector space $T\Omega_t$. We denote by $\mathbf{F}_m$ the linear map $T_{\Psi(m)}\mathcal{B} \to T_m\Omega_t$ (the deformation gradient, see [39]) defined by

$$\mathbf{F}_m := (T_m \Psi_t)^{-1}.$$
Remark 4.1. The deformation gradient **F** corresponds to the right inverse $(T\Psi)_{\mathbf{N}^{\perp}}^{-1}$, which is characterized by

$$T\Psi \cdot (T\Psi)_{\mathbf{N}^{\perp}}^{-1} = \mathrm{id}_{V}, \quad \text{and} \quad \mathbf{N}^{\flat} \cdot (T\Psi)_{\mathbf{N}^{\perp}}^{-1} = 0.$$

Similarly, the right inverse $(T\Psi)_{\mathbf{U}^{\perp}}^{-1}$ of the linear tangent map $T\Psi$ is the unique solution of

$$T\Psi \cdot (T\Psi)_{\mathbf{U}^{\perp}}^{-1} = \mathrm{id}_{V}, \quad \text{and} \quad \mathbf{U}^{\flat} \cdot (T\Psi)_{\mathbf{U}^{\perp}}^{-1} = 0.$$

The relation between these two right inverses is then given by

$$(T\Psi)_{\mathbf{U}^{\perp}}^{-1} = \mathbf{U} \otimes (\mathbf{U}^{\flat} \cdot \mathbf{F}) + \mathbf{F}.$$

Using the fact that

$$T\Psi \cdot \mathbf{U} = \gamma \left( T\Psi \cdot \mathbf{N} + \frac{1}{c} T\Psi \cdot \mathbf{u} \right) = \gamma \left( T\Psi \cdot \mathbf{N} + \frac{1}{c} T\Psi_t \cdot \mathbf{u} \right) = 0,$$

since $\mathbf{u}$ is tangent to $\Omega_t$, we get

$$T\Psi \cdot \mathbf{N} = -\frac{1}{c} \mathbf{F}^{-1} \mathbf{u}.$$

The orthogonal decomposition of $T\Psi$ relative to $\mathbf{N}$ recast thus as

$$T\Psi = \frac{1}{c}(\mathbf{F}^{-1}\mathbf{u}) \otimes \mathbf{N}^{\flat} + \mathbf{F}^{-1}, \quad \mathbf{F}^{-1}\mathbf{N} = 0.$$

The column vectors $\mathbf{F}_J := \mathbf{F} \cdot \partial/\partial X^J$ of $\mathbf{F}$ in the canonical basis of $\mathbb{R}^3$ are characterized by the properties:

$$(4.2) \qquad \mathbf{N}^{\flat}(\mathbf{F}_J) = 0 \quad \text{and} \quad d\Psi_t^I(\mathbf{F}_J) = \delta_J^I.$$

We shall define the *observer frame* as

$$\mathcal{F}_{\mathrm{obs}} := (\mathbf{N}, \mathbf{F}_1, \mathbf{F}_2, \mathbf{F}_3).$$

Beware that the vector fields in this frame are *not* general covariant of $g$ and $\Psi$.

Remark 4.2. The frame $\mathcal{F}_{\mathrm{rest}}$ introduced in theorem 2.6 is linked to the matter and is thus the *rest frame* of the matter (meaning that the spatial velocity $\mathbf{u} = 0$ and the Lorentz factor $\gamma = 1$ in this frame). In General Relativity, any frame

$$\mathcal{F} := (\mathbf{N}, e_i), \quad \text{with} \quad g(\mathbf{N}, e_i) = 0,$$

respecting the orthogonal $(3+1)$-decomposition induced by a time function is called an *Eulerian frame* [31]. Our observer frame $\mathcal{F}_{\mathrm{obs}}$ is thus a particular choice of an Eulerian frame.

**4.2. Rewriting the invariants in the observer frame.** Rewriting the vector valued invariants $\mathbf{A}$, $\mathbf{L}$ and $\mathbf{H}$ using the components of the Hessian $\mathbf{Hess}^g \Psi^I$ in the observer frame $(\mathbf{N}, \mathbf{F}_1, \mathbf{F}_2, \mathbf{F}_3)$ will give more mechanistic expressions than the initial expressions (3.3), (3.5) and (3.7). These expressions will be used to compute their classical limit, and to recover the invariants of classical second gradient hyperelasticity. However, in order to get a compact statement and avoid enlarging the formulas, we will first keep the vector field $\mathbf{U}$ and express the invariants in the frame $(\mathbf{U}, \mathbf{F}_1, \mathbf{F}_2, \mathbf{F}_3)$ rather than $(\mathbf{N}, \mathbf{F}_1, \mathbf{F}_2, \mathbf{F}_3)$, as illustrated in the following result.

**Lemma 4.3.** We have

$$\left\{
\begin{aligned}
A^I &= d\Psi^I \cdot \nabla_{c\mathbf{U}}^g (c\mathbf{U}), \\
L_J^I &= d\Psi^I \cdot \nabla_{\mathbf{F}_J}^g (c\mathbf{U}) + \frac{1}{c} A^I \mathbf{U}^{\flat} (\mathbf{F}_J), \\
H_{JK}^I &= d\Psi^I \cdot \nabla_{\mathbf{F}_K}^g (\mathbf{F}_J) + \frac{1}{c} \left( L_J^I \mathbf{U}^{\flat} (\mathbf{F}_K) + L_K^I \mathbf{U}^{\flat} (\mathbf{F}_J) \right) + \frac{1}{c^2} A^I \mathbf{U}^{\flat} (\mathbf{F}_J) \mathbf{U}^{\flat} (\mathbf{F}_K),
\end{aligned}
\right.$$
or, in tensorial notation

$$
\left\{
\begin{aligned}
\mathbf{A} &= T\Psi \cdot \nabla_{cU}^g (c\mathbf{U}), \\
\mathbf{L} &= T\Psi \cdot (\nabla^g c\mathbf{U}) \cdot \mathbf{F} + \frac{1}{c} \mathbf{A} \otimes (\mathbf{U}^\flat \mathbf{F}), \\
\mathbf{H} &= (T\Psi \cdot \nabla_{\mathbf{F}_K}^g (\mathbf{F}_J)) \otimes dX^J \otimes dX^K + \frac{2}{c} (\mathbf{L} \otimes (\mathbf{U}^\flat \mathbf{F}))^{(23)} + \frac{1}{c^2} \mathbf{A} \otimes (\mathbf{U}^\flat \mathbf{F}) \otimes (\mathbf{U}^\flat \mathbf{F}),
\end{aligned}
\right.
$$

where $\langle \cdot \rangle^{(23)}$ means the symmetrization over the second and third factors.

*Proof.* From (2.9) and (4.2), and the fact that $d\Psi^I \cdot \mathbf{F} = dX^I$, one deduces that

$$ (4.3) \qquad E_J = F_J + U^b(F_J)U, $$

and one has

$$ d\Psi^I \cdot \mathbf{U} = 0 \quad \text{and} \quad d\Psi^I \cdot E_J = d\Psi^I \cdot F_J = \delta_J^I. $$

Now, using the Leibniz rule

$$ (\mathbf{Hess}^g \Psi^I)(\mathbf{X}, \mathbf{Y}) = \nabla_{\mathbf{X}}^g d\Psi^I \cdot \mathbf{Y} = d(d\Psi^I \cdot \mathbf{X}) \cdot \mathbf{Y} - d\Psi^I \cdot \nabla_{\mathbf{X}}^g \mathbf{Y}, $$

and the fact that the Hessian is symmetric (because the Levi-Civita connection is torsionless), one gets

$$ (\mathbf{Hess}^g \Psi^I)(c\mathbf{U}, c\mathbf{U}) = -d\Psi^I \cdot \nabla_{c\mathbf{U}}^g (c\mathbf{U}), $$

$$ (\mathbf{Hess}^g \Psi^I)(c\mathbf{U}, \mathbf{F}_J) = (\mathbf{Hess}^g \Psi^I)(\mathbf{F}_J, c\mathbf{U}) = -d\Psi^I \cdot \nabla_{c\mathbf{U}}^g (\mathbf{F}_J) = -d\Psi^I \cdot \nabla_{\mathbf{F}_J}^g (c\mathbf{U}), $$

$$ (\mathbf{Hess}^g \Psi^I)(\mathbf{F}_J, \mathbf{F}_K) = (\mathbf{Hess}^g \Psi^I)(\mathbf{F}_K, \mathbf{F}_J) = -d\Psi^I \cdot \nabla_{\mathbf{F}_J}^g (\mathbf{F}_K) = -d\Psi^I \cdot \nabla_{\mathbf{F}_K}^g (\mathbf{F}_J). $$

We get therefore, using (4.3) and the bilinearity of the Hessian.

* Firstly:

$$ A^I = -c^2 (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{U}) = d\Psi^I \cdot \nabla_{c\mathbf{U}}^g (c\mathbf{U}), $$

* Secondly:

$$
\begin{align*}
L_J^I &= -c (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{E}_J) \\
&= -c (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{F}_J + \mathbf{U}^b(\mathbf{F}_J)\mathbf{U}) \\
&= -c (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{F}_J) - c (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{U}) \mathbf{U}^b(\mathbf{F}_J) \\
&= d\Psi^I \cdot \nabla_{\mathbf{F}_J}^g (c\mathbf{U}) + \frac{1}{c} A^I \mathbf{U}^b(\mathbf{F}_J).
\end{align*}
$$

* Thirdly:

$$
\begin{align*}
H_{JK}^I &= -(\mathbf{Hess}^g \Psi^I)(\mathbf{E}_J, \mathbf{E}_K) \\
&= -(\mathbf{Hess}^g \Psi^I)(\mathbf{F}_J + \mathbf{U}^b(\mathbf{F}_J)\mathbf{U}, \mathbf{F}_K + \mathbf{U}^b(\mathbf{F}_K)\mathbf{U}) \\
&= -(\mathbf{Hess}^g \Psi^I)(\mathbf{F}_J, \mathbf{F}_K) - (\mathbf{Hess}^g \Psi^I)(\mathbf{F}_J, \mathbf{U}) \mathbf{U}^b(\mathbf{F}_K) \\
&\quad - (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{F}_K) \mathbf{U}^b(\mathbf{F}_J) - (\mathbf{Hess}^g \Psi^I)(\mathbf{U}, \mathbf{U}) \mathbf{U}^b(\mathbf{F}_J) \mathbf{U}^b(\mathbf{F}_K) \\
&= d\Psi^I \cdot \nabla_{\mathbf{F}_J}^g (\mathbf{F}_K) + \frac{1}{c} L_J^I \mathbf{U}^b(\mathbf{F}_K) + \frac{1}{c} L_K^I \mathbf{U}^b(\mathbf{F}_J) + \frac{1}{c^2} A^I \mathbf{U}^b(\mathbf{F}_J) \mathbf{U}^b(\mathbf{F}_K).
\end{align*}
$$

□

**4.3. Second Gradient Hyperelasticity in Minkowski spacetime.** We will now specify the general theory in a familiar simple setting. Since the presence of the studied matter in the laboratory does not affect (much) the Universe metric $g$ compared to the presence of the Sun, of the Earth and of other planets (passive matter assumption), we can consider that $g$ is now fixed and choose, among the numerous spacetimes encountered in General Relativity and available in the catalog [49], those describing solutions of Einstein equations in the vacuum. In this paper, we have chosen to work in the Minkowski spacetime $(\mathcal{M}, \eta)$, the framework of
Special Relativity. It is a flat pseudo-Riemannian manifold, diffeomorphic to $\mathbb{R}^4$ with canonical coordinates $(x^0 = ct, x^i)$ and endowed with the static metric

$$ \eta = -c^2 \, d t^2 + \mathfrak{g}, \quad \text{where} \quad \mathfrak{g} = \delta_{ij} \, d x^i \, d x^j. $$

The time function is here $\hat{t}(m) = t = x^0/c$ and accordingly, one has

$$ \boldsymbol{N} = \frac{1}{c} \frac{\partial}{\partial t}, \quad \text{and} \quad \boldsymbol{N}^{\flat} = -c \, d t. $$

Note that in this simple case $\nabla^{\eta} \boldsymbol{N} = 0$. As a consequence, for any vector field $\boldsymbol{X}$ tangent to $\mathcal{E}_t$, one has

$$ 0 = d g(\boldsymbol{X}, \boldsymbol{N}) = g(\nabla^{\eta} \boldsymbol{X}, \boldsymbol{N}) + g(\nabla^{\eta} \boldsymbol{N}, \boldsymbol{X}), $$

and thus

$$ g(\nabla^{\eta} \boldsymbol{X}, \boldsymbol{N}) = -g(\nabla^{\eta} \boldsymbol{N}, \boldsymbol{X}) = 0. $$

For such vector fields, at a point of $\Omega_t$, one has therefore

$$ T \Psi \cdot \nabla^{\eta} \boldsymbol{X} = T \Psi_t \cdot \nabla^{\eta} \boldsymbol{X} = \boldsymbol{F}^{-1} \cdot \nabla^{\eta} \boldsymbol{X}, $$

where $\boldsymbol{F} := T \Psi_t^{-1} : T \mathcal{B} \to T \Omega_t$. If $\boldsymbol{X}$ and $\boldsymbol{Y}$ are both tangent to $\mathcal{E}_t$, one has furthermore

$$ \nabla_{\boldsymbol{X}}^{\eta} \boldsymbol{Y} = \nabla_{\boldsymbol{Y}}^{\mathfrak{g}} \boldsymbol{Y}, $$

i.e. the hypersurface $\mathcal{E}_t$ of $(\mathcal{M}, \mathfrak{g})$ has no extrinsic curvature. Here, $(\mathcal{E}_t, \mathfrak{g})$ is a three dimensional Euclidean affine space. Finally, the quadrivelocity and Lorentz factor have for expressions

$$ c \boldsymbol{U} = \gamma \left( \frac{\partial}{\partial t} + \boldsymbol{u} \right), \quad \gamma = \frac{1}{\sqrt{1 - \frac{\| \boldsymbol{u} \|_{\mathfrak{g}}^2}{c^2}}}. $$

and one has

$$ (4.4) \qquad \boldsymbol{U}^{\flat} \boldsymbol{F} = \frac{\gamma}{c} \boldsymbol{u}^{\flat} \boldsymbol{F}. $$

**Remark 4.4.** One can observe that $\gamma^{-1} \nabla_c^{\eta} \boldsymbol{U} \boldsymbol{X}$ corresponds here to the (relativistic) particle derivative, which we denote with a dot,

$$ \dot{\boldsymbol{X}} := \gamma^{-1} \nabla_c^{\eta} \boldsymbol{U} \boldsymbol{X} = \partial_t \boldsymbol{X} + \nabla_{\boldsymbol{u}}^{\mathfrak{g}} \boldsymbol{X}, $$

and that if $\boldsymbol{X}$ is tangent to $\mathcal{E}_t$, then so is $\dot{\boldsymbol{X}}$.

Since the curvature tensor vanishes, the invariants $\boldsymbol{M}_2, \boldsymbol{M}_3$ and $\boldsymbol{M}_4$ vanish and we need only to explicit the remaining vector valued relativistic invariants $\boldsymbol{K}, \boldsymbol{A}, \boldsymbol{L}$ and $\boldsymbol{H}$.

**Theorem 4.5.** In the Minkowski spacetime:

* The conformation $\boldsymbol{K} = (K^{IJ})$ is given by

$$ \boldsymbol{K} = \boldsymbol{F}^{-1} \left( \mathfrak{g}^{-1} - \frac{1}{c^2} \boldsymbol{u} \otimes \boldsymbol{u} \right) \boldsymbol{F}^{-\star}, $$

which corresponds to the formula derived in [61].

* The material relativistic acceleration $\boldsymbol{A} = (A^I)$ is given by

$$ \boldsymbol{A} = \boldsymbol{F}^{-1} \boldsymbol{a}, \quad \text{where} \quad \boldsymbol{a} := \gamma^2 \dot{\boldsymbol{u}} = \gamma^2 (\partial_t \boldsymbol{u} + \nabla_{\boldsymbol{u}}^{\mathfrak{g}} \boldsymbol{u}). $$

Observe that $\boldsymbol{A}$ is the pullback on the body $\mathcal{B}$ of the spatial relativistic acceleration $\boldsymbol{a}$, defined on $\Omega_t$.

* The material relativistic velocity gradient $\boldsymbol{L} = (L^{IJ})$ is given by

$$ \boldsymbol{L} = \boldsymbol{F}^{-1} \cdot \left( \gamma \nabla_{\boldsymbol{g}} \boldsymbol{u} + \frac{\gamma}{c^2} \boldsymbol{a} \otimes \boldsymbol{u}^{\flat} \right) \cdot \boldsymbol{F} $$

Observe that $\boldsymbol{L}$ is the pullback on the body $\mathcal{B}$ of the spatial mixed tensor

$$ \gamma \nabla_{\boldsymbol{g}} \boldsymbol{u} + \frac{\gamma}{c^2} \boldsymbol{a} \otimes \boldsymbol{u}^{\flat}, $$

defined on $\Omega_t$.
• The material relativistic strain gradient $\mathbf{H} = (H_{JK}^I)$ is given by

$$ (4.5) \qquad \mathbf{H} = \mathbf{F}^{-1} \cdot \nabla^g \mathbf{F} \cdot \mathbf{F} + \frac{2\gamma}{c^2} (\mathbf{L} \otimes (\mathbf{u}^b \mathbf{F}))^{(23)} + \frac{\gamma^2}{c^4} \mathbf{A} \otimes (\mathbf{u}^b \mathbf{F}) \otimes (\mathbf{u}^b \mathbf{F}). $$

where $\mathbf{u}^{(23)}$ means the symmetrization over the second and third factors, and

$$ (\nabla^g \mathbf{F})_{Jk}^i := \frac{\partial F_J^i}{\partial x^k}. $$

Remark 4.6. In Classical Continuum Mechanics, one starts with an embedding

$$ \mathbf{p}: \mathcal{B} \to \mathcal{E}, \qquad \mathbf{X} \mapsto \mathbf{x} = \mathbf{p}(\mathbf{X}), $$

possibly parameterized by time, and introduces the *deformation gradient*

$$ \mathbf{F}_p(\mathbf{X}) := T_{\mathbf{X}p}: T_{\mathbf{X}}\mathcal{B} \to T_{p(\mathbf{X})}\mathcal{E}. $$

In our setting, if we assume that the restriction $\Psi_t: \Omega_t \to \mathcal{B}$ is a diffeomorphism and we set $\mathbf{p} := \Psi_t^{-1}$, then, we have $\mathbf{F}_p(\mathbf{X}) = \mathbf{F}(\mathbf{p}(\mathbf{X}))$ and

$$ \frac{\partial(\mathbf{F}_p)^i_J}{\partial X^K} = \frac{\partial F_J^i}{\partial x^l} (\mathbf{F}_p)^l_K, $$

or, in tensorial notation,

$$ \nabla \mathbf{F}_p = \nabla^g \mathbf{F} \cdot \mathbf{F}_p = \nabla^g \mathbf{F} \cdot \mathbf{F} \circ \mathbf{p}. $$

This means that the left inner inside term of (4.5) can be rewritten

$$ \mathbf{F}^{-1} \cdot \nabla^g \mathbf{F} \cdot \mathbf{F} = (\mathbf{F}_p^{-1} \cdot \nabla \mathbf{F}_p) \circ \mathbf{p}^{-1}, $$

so as to display the standard deformation gradient variable $\mathbf{F}_p^{-1} \cdot \nabla \mathbf{F}_p$ on $\mathcal{B}$ of Classical Continuum Mechanics.

*Proof.* We have first the orthogonal decompositions relative to $\mathbf{N}$

$$ g^{-1} = \mathfrak{g}^{-1} - \mathbf{N} \otimes \mathbf{N}, \quad T\mathfrak{g} = \frac{1}{c}(\mathbf{F}^{-1}\mathbf{u}) \otimes \mathbf{N}^b + \mathbf{F}^{-1}, \quad T\mathfrak{g}^* = \frac{1}{c}\mathbf{N}^b \otimes (\mathbf{F}^{-1}\mathbf{u}) + \mathbf{F}^{-*}. $$

We get thus

$$ \mathbf{K} = T\mathfrak{g} \cdot g^{-1} \cdot T\mathfrak{g}^* = \mathbf{F}^{-1}\mathfrak{g}^{-1}\mathbf{F}^{-*} - \frac{1}{c^2}(\mathbf{F}^{-1}\mathbf{u}) \otimes (\mathbf{F}^{-1}\mathbf{u}) = \mathbf{F}^{-1}\left(\mathfrak{g}^{-1} - \frac{1}{c^2}\mathbf{u} \otimes \mathbf{u}\right)\mathbf{F}^{-*}. $$

We will now particularize the formulas given in lemma 4.3 when $g = \eta$. In that case, we have

$$ \nabla^\eta \mathbf{N} = 0 \quad \text{and} \quad c\mathbf{U} = \gamma(c\mathbf{N} + \mathbf{u}), \quad \text{with} \quad \mathbf{N} = \frac{1}{c} \frac{\partial}{\partial t}. $$

Note also that since $\nabla^\eta \mathbf{N} = 0$ and $T\mathfrak{g} \cdot \mathbf{U} = 0$, we get that

$$ \begin{aligned} \nabla_X^\eta (c\mathbf{U}) &= \nabla_X^\eta (\gamma(c\mathbf{N} + \mathbf{u})) \\ &= (d\gamma \cdot \mathbf{X})(c\mathbf{N} + \mathbf{u}) + \gamma \nabla_X^\eta (c\mathbf{N} + \mathbf{u}) \\ &= \frac{1}{\gamma}(d\gamma \cdot \mathbf{X})c\mathbf{U} + \gamma \nabla_X^\eta \mathbf{u}, \end{aligned} $$

for any vector field $\mathbf{X}$. Hence

$$ T\mathfrak{g} \cdot \nabla^\eta (c\mathbf{U}) = \gamma T\mathfrak{g} \cdot \nabla^\eta \mathbf{u}. $$

We get therefore, using remark 4.4,

$$ \mathbf{A} = T\mathfrak{g} \cdot \nabla_{c\mathbf{U}}^g (c\mathbf{U}) = \gamma T\mathfrak{g} \cdot \nabla_{c\mathbf{U}}^\eta \mathbf{u} = \gamma^2 T\mathfrak{g} \cdot \nabla_{c\mathbf{U}}^\eta \dot{\mathbf{u}} = \gamma^2 \mathbf{F}^{-1} \dot{\mathbf{u}} = \mathbf{F}^{-1} \mathbf{a}, $$
then, using (4.4),

$$
\begin{align*}
\mathbf{L} &= T\Psi \cdot (\nabla^g c\mathbf{U}) \cdot \mathbf{F} + \frac{1}{c} \mathbf{A} \otimes (\mathbf{U}^b \mathbf{F}) \\
&= \gamma T\Psi \cdot \nabla^\eta \mathbf{u} \cdot \mathbf{F} + \frac{\gamma}{c^2} (\mathbf{F}^{-1}\mathbf{a}) \otimes (\mathbf{u}^b \mathbf{F}) \\
&= \gamma \mathbf{F}^{-1} \cdot \nabla^g \mathbf{u} \cdot \mathbf{F} + \frac{\gamma}{c^2} \mathbf{F}^{-1} \cdot (\mathbf{a} \otimes \mathbf{u}^b) \cdot \mathbf{F} \\
&= \mathbf{F}^{-1} \cdot \left( \gamma \nabla^g \mathbf{u} + \frac{\gamma}{c^2} \mathbf{a} \otimes \mathbf{u}^b \right) \cdot \mathbf{F}
\end{align*}
$$

and, using (4.4),

$$
\mathbf{H} = (T\Psi \cdot \nabla_{\mathbf{F}_J}^{\eta}(\mathbf{F}_K)) \otimes dX^J \otimes dX^K + \frac{2\gamma}{c^2} (\mathbf{L} \otimes (\mathbf{u}^b \mathbf{F}))^{(23)} + \frac{\gamma^2}{c^4} \mathbf{A} \otimes (\mathbf{u}^b \mathbf{F}) \otimes (\mathbf{u}^b \mathbf{F}).
$$

But

$$
T\Psi \cdot \nabla_{\mathbf{F}_J}^{\eta}(\mathbf{F}_K) = T\Psi_t \cdot \nabla_{\mathbf{F}_J}^{g}(\mathbf{F}_K) = \mathbf{F}^{-1} \cdot \nabla_{\mathbf{F}_J}^{g}(\mathbf{F}_K),
$$

where we have defined $\mathbf{F}_J$ as $\mathbf{F} \cdot \partial/\partial X^J$ and where $\mathbf{F}(m) = (T_m \Psi_t)^{-1}$ for each $m \in \Omega_t$. We can thus write in components

$$
(\mathbf{F}_J)^i = F_J^i, \quad \text{and} \quad (\nabla_{\mathbf{F}_J}^g (\mathbf{F}_K))^i = \frac{\partial (\mathbf{F}_K)^i}{\partial x^l} (\mathbf{F}_J)^l = \frac{\partial F_K^i}{\partial x^l} F_J^l.
$$

We get thus

$$
(\mathbf{F}^{-1} \nabla_{\mathbf{F}_J}^{g} \mathbf{F}_K)^I = (\mathbf{F}^{-1} \nabla_{\mathbf{F}_K}^{g} \mathbf{F}_J)^I = (\mathbf{F}^{-1})^I{}_i \frac{\partial F_J^i}{\partial x^l} F_K^l = (\mathbf{F}^{-1})^I{}_i (\nabla^g \mathbf{F})^i_{Jl} F_K^l,
$$

if we set

$$
(\nabla^g \mathbf{F})^i_{Jk} := \frac{\partial F_J^i}{\partial x^k}.
$$

We have finally

$$
\mathbf{H} = \mathbf{F}^{-1} \cdot \nabla^g \mathbf{F} \cdot \mathbf{F} + \frac{2\gamma}{c^2} (\mathbf{L} \otimes (\mathbf{u}^b \mathbf{F}))^{(23)} + \frac{\gamma^2}{c^4} \mathbf{A} \otimes (\mathbf{u}^b \mathbf{F}) \otimes (\mathbf{u}^b \mathbf{F}). \quad \square
$$

5. CLASSICAL LIMIT OF GRADIENT RELATIVISTIC HYPERELASTICITY

5.1. **Classical limit of the relativistic invariants.** As the speed of light $c$ goes to infinity, the conformation tends towards the inverse of the Cauchy–Green tensor $\mathbf{C} := \mathbf{F}^*\mathbf{g}\mathbf{F}$ of Classical Continuum Mechanics [59, 43],

$$
\lim_{c \to \infty} \mathbf{K} = \mathbf{C}^{-1}.
$$

The vector **A** depends on the speed of light *c* through the Lorentz factor $\gamma$. At the classical
limit, since $\lim_{c \to \infty} \gamma = 1$, we have then

$$
\lim_{c \to \infty} \mathbf{A} = \mathbf{F}^{-1} \mathbf{a}, \quad \mathbf{a} = \dot{\mathbf{u}} = \partial_t \mathbf{u} + \nabla_u^g \mathbf{u},
$$

meaning that **A**, interpreted as the pullback on $\mathcal{B}$ of the Eulerian acceleration **a**, is the material acceleration of Classical Continuum Mechanics.

At the classical limit

$$
\lim_{c \to \infty} \mathbf{L} = \mathbf{F}^{-1} (\nabla^g \mathbf{u}) \mathbf{F} = \mathbf{F}^{-1} \mathbf{l} \mathbf{F}, \quad \mathbf{l} := \nabla^g \mathbf{u},
$$

This tensor is interpreted as the pullback on $\mathcal{B}$ of the so-called Eulerian velocity gradient $\boldsymbol{l} := \nabla^g \boldsymbol{u}$ (a mixed second order tensor on $\Omega_t$). It is the material (on $\mathcal{B}$) velocity gradient of Classical Continuum Mechanics. The infinite speed of light limit of $\mathbf{L}$ is thus built from the sum of the (symmetric) strain rate tensor $\mathbf{d} := \boldsymbol{l}^s = (\nabla^g \boldsymbol{u})^s$, which is known to be objective, and of the (skew-symmetric) spin tensor $\mathbf{w} := \boldsymbol{l}^a = (\nabla^g \boldsymbol{u})^a$, which is known to be non-objective.

Lastly, the contravariant third order tensor **H** has for limit at infinite speed of light

$$
\lim_{c \to \infty} \mathbf{H} \circ \mathbf{p} = \mathbf{F}_{\mathrm{p}}^{-1} \nabla \mathbf{F}_{\mathrm{p}}, \quad \mathbf{F}_{\mathrm{p}} = T \mathrm{p},
$$
where $p: \mathcal{B} \to \mathcal{E}$ is the embedding of the body in the Euclidean space (see Remark 4.6). The secondary variable $\mathbf{F}_p^{-1}\nabla\mathbf{F}_p$, obtained as such an infinite speed of light limiting case, is the fundamental variable (objective, see discussion below) of Toupin and Mindlin finite strain second gradient theory [23]. It is equal to $\mathbf{F}_\phi^{-1}\nabla\mathbf{F}_\phi$ when the body $\mathcal{B}$ is identified with the reference configuration $\Omega_0$, $\phi: \Omega_0 \to \Omega_t$ is the deformation and $\mathbf{F}_\phi = T\phi$ is the deformation gradient.

**Remark 5.1.** The alternative (contravariant) invariants $\mathbf{L}^{\sharp\kappa^{-1}} = \mathbf{L} \cdot \mathbf{K}$ and $\mathbf{H}^{\sharp\kappa^{-1}}$ introduced in **Remark 3.7** have as classical limits (at $c \to \infty$)

$$ \lim_{c \to \infty} \mathbf{L}^{\sharp\kappa^{-1}} = \mathbf{F}^{-1}(\nabla^g \boldsymbol{u}) \mathbf{F} \cdot \mathbf{C}^{-1} \quad \text{and} \quad \lim_{c \to \infty} \mathbf{H}^{\sharp\kappa^{-1}} \circ \boldsymbol{p} = (\mathbf{F}_p^{-1}\nabla\mathbf{F}_p)^{\sharp\mathbf{C}^{-1}}, $$

the latter of components

$$ (\mathbf{F}_p^{-1}\nabla\mathbf{F}_p)^I_{PQ}(\mathbf{C}^{-1})^{PI}(\mathbf{C}^{-1})^{QJ}. $$

**5.2. From General Covariance to Objectivity.** Starting from General Relativity and fixing the metric to the flat Minkowski metric $\eta$ on $\mathcal{M} \simeq \mathbb{R}^4$ corresponds to the framework of Special Relativity. It is a symmetry breaking because in this setting, only diffeomorphisms $\varphi$ which fix $\eta$, meaning that $\varphi^*\eta = \eta$, are allowed. These diffeomorphisms are the isometries of the Minkowski space and constitute the Poincaré group, a finite dimensional Lie group of dimension 10. In Classical Continuum Mechanics, the Universe $\mathcal{M} \simeq \mathbb{R}^4$ is no longer a Lorentzian manifold. It is equipped with a Galilean structure, i.e., a pair $(\kappa, \theta)$, where $\kappa$ is a symmetric second-order contravariant tensor of signature $(0, +, +, +)$ (the classical spatial cometric) and $\theta$ is a 1-form which spans the kernel of $\kappa$ (the clock) [39]. In the present case, this Galilean structure $(\kappa, \theta)$ is obtained as the limit of the 1-parameter family of smooth Lorentz metrics

$$ \eta_c^{-1} = -c^{-2}\partial_t^2 + \partial_x^2 + \partial_y^2 + \partial_z^2, $$

when $c \to \infty$, and where $\theta = dt$. Diffeomorphisms $\varphi$ which preserve the Galilean structure

$$ \varphi^*\kappa = \kappa, \quad \text{and} \quad \varphi^*\theta = \theta, $$

are called *Gallileomorphisms*. In the canonical coordinate system of $\mathbb{R}^4$, they are given by

$$ \bar{\mathbf{x}} = \mathbf{Q}(t)\mathbf{x} + \mathbf{b}(t), \quad \bar{t} = t + c, $$

where $\mathbf{x} = (x, y, z)$, $\mathbf{Q}(t)$ is a rotation and $\mathbf{b}(t)$ is a vector, both depending on time. These transformations are the “change of observer” of Classical Mechanics. Those quantities which are invariant under these transformations are called *objective* or *frame independent* [40]. As a result of our study, it appears that even if all the quantities of Corollary 3.6 are invariant under general (orientation-preserving) diffeomorphisms, their limit may or may not be objective. The reason for that is that some quantities like the conformation $\mathbf{K}$ or the material relativistic strain gradient $\mathbf{H}$, defined a priori for Lorentzian metrics $g$, extend continuously on the boundary of this set (that is, to degenerate Lorentzian cometrics) and thus to Galilean structures. In that case the General Covariance of $\mathbf{K}$ and $\mathbf{H}$ leads to the objectivity of their limits $\mathbf{C}^{-1}$ and $\mathbf{F}_p^{-1}\nabla\mathbf{F}_p$, by restricting to symmetries of the Galilean structure. However, this is not the case for other invariants like the material relativistic acceleration $\mathbf{A}$ and the material relativistic velocity gradient $\mathbf{L}$.

Such results in fact clarify the criticism levelled at the principle of objectivity, which is sometimes considered too restrictive, for instance for viscous fluids [17, 5, 50, 63]: indeed, the general covariant secondary variables which do not have objective limits are in fact first time-gradient and first-spatial gradient (for $\mathbf{A}$ and $\mathbf{L}$). All general covariant first-order ($\mathbf{K}$) and second-order ($\mathbf{H}$) spatial gradient variables authorized by Theorem 3.3 for solids second-gradient hyperelasticity have objective limits at $c \to \infty$.

## 6. CONCLUSION

We have extended Souriau's theory of Relativistic first gradient Hyperelasticity to a second gradient theory. We have shown that, assuming General Covariance for the Lagrangian density $L$, it is in fact a function of the following diffeomorphisms invariants: the matter field $\Psi$ itself, Souriau's conformation $\mathbf{K}$ and additional three-dimensional tensorial valued invariants, namely
* the material relativistic acceleration (vector **A**),

* the material relativistic velocity gradient (second order tensor **L**),

* the material relativistic strain gradient (third order tensor **H**),

* and relativistic gravitation/matter coupling tensors (second, third and fourth order tensors **M**<sub>p</sub>), built from the curvature tensor **R**<sup>g</sup>.

They play the role of Souriau's conformation, but for a relativistic second gradient theory. The Galilean limits of these invariants have been calculated in the flat Minkowski spacetime of Special Relativity. Our results show that the 3D Classical Continuum Mechanics second gradient theory can be derived as limiting case $c \to \infty$ of such a relativistic theory. Some secondary variables converge in the Galilean limit to objective quantities (such as the fundamental variable $\mathbf{F}_{\phi}^{-1}\nabla\mathbf{F}_{\phi}$ of finite strain second gradient theory, modeling solids), other secondary variables to non-objective quantities (modeling fluids). The present work thus contributes to the debate on the criticism of the Principle of Objectivity, sometimes considered too restrictive, particularly in fluid mechanics [17, 5, 50, 63].

## APPENDIX A. COVARIANT DERIVATIVES, TORSION, CURVATURE AND HESSIAN TENSORS

**A.1. Covariant derivatives.** A covariant derivative $\nabla$ on $\mathcal{M}$ is an operator which extends the usual derivative of scalar fields

$$ \nabla_X f := d f \cdot X, \quad f \in C^\infty(\mathcal{M}), \quad \mathbf{X} \in \mathcal{X}^1(\mathcal{M}), $$

to tensor fields. It is characterized by the following two properties:

$$ \nabla_{fX} Y = f \nabla_X Y, \quad \nabla_X(fY) = \nabla_X(f)Y + f \nabla_X(Y), $$

for all functions $f \in C^\infty(\mathcal{M})$ and all vector fields $\mathbf{X}, \mathbf{Y}$ on $\mathcal{M}$. Using Leibniz rule, it extends naturally to covector fields $\alpha \in \Omega^1(\mathcal{M})$, via the formula

$$ (\nabla_X \alpha)(Y) := \nabla_X(\alpha(Y)) - \alpha(\nabla_X Y), $$

and more generally to all tensor fields. We admit that if **t** is a tensor field of order *r*, then $\nabla**t$ is also a tensor field, of order (*r* + 1).

**Example A.1.** For a metric *g*, Leibniz rule yields

$$ (\nabla_Z g)(\mathbf{X}, \mathbf{Y}) := d(g(\mathbf{X}, \mathbf{Y})) \cdot \mathbf{Z} - g(\nabla_Z \mathbf{X}, \mathbf{Y}) - g(\mathbf{X}, \nabla_Z \mathbf{Y}). $$

Given a local coordinates system ($x^\mu$), a covariant derivative $\nabla$ is determined by its Christoffel coefficients (of the second kind)

$$ \Gamma^\kappa_{\lambda\mu} := d x^\kappa (\nabla_{\partial_\mu} \partial_\lambda) $$

Note that despite the tensorial notation, these are not the components of a tensor.

**Definition A.2.** Denote by $[\mathbf{X}, \mathbf{Y}] \in \mathcal{X}^1(\mathcal{M})$ the Lie bracket of two vector fields **X** and **Y** on $\mathcal{M}$. Given a covariant derivative $\nabla$, one defines the following tensor fields.

(1) the *torsion tensor*

$$ \mathbf{T}^{\nabla}(\mathbf{X}, \mathbf{Y}) := \nabla_{\mathbf{X}}\mathbf{Y} - \nabla_{\mathbf{Y}}\mathbf{X} - [\mathbf{X}, \mathbf{Y}], $$

with components

$$ T^\kappa_{\lambda\mu} = \Gamma^\kappa_{\lambda\mu} - \Gamma^\kappa_{\mu\lambda}; $$

(2) the *Riemann curvature tensor*

$$ \mathbf{R}^{\nabla}(\mathbf{X}, \mathbf{Y})\mathbf{Z} := \nabla_{\mathbf{X}}\nabla_{\mathbf{Y}}\mathbf{Z} - \nabla_{\mathbf{Y}}\nabla_{\mathbf{X}}\mathbf{Z} - \nabla_{[\mathbf{X}, \mathbf{Y}]}\mathbf{Z}, $$

with components

$$ R^\kappa_{\lambda\mu\nu} = \partial_\mu \Gamma^\kappa_{\nu\lambda} - \partial_\nu \Gamma^\kappa_{\lambda\mu} + \Gamma^\kappa_{\mu\rho} \Gamma^\rho_{\nu\lambda} - \Gamma^\kappa_{\nu\rho} \Gamma^\rho_{\lambda\mu}; $$
(3) the Hessian tensor of a scalar field $f$

$$ \mathbf{Hess}^{\nabla} f := \nabla \nabla f = \nabla(\mathrm{d} f), $$

with components

$$ (\mathbf{Hess}^{\nabla} f)_{\mu\nu} = \partial_{\mu}\partial_{\nu}f - \partial_{\kappa}f \Gamma^{\kappa}_{\nu\mu}. $$

A covariant derivative $\nabla$ is said to be symmetric if its torsion tensor $\mathbf{T}^{\nabla}$ vanishes or, in other words, if $\Gamma^{\kappa}_{\lambda\mu}$ is symmetric in the lower indices. In that case, the Hessian $\mathbf{Hess}^{\nabla} f$ of any scalar field is also symmetric.

Recall that the Universe $(\mathcal{M}, g)$ is assumed to be a pseudo-Riemannian manifold where $g$ is Lorentzian metric of signature $(- + ++)$. Such a manifold is endowed with a natural covariant derivative $\nabla^g$, the Levi-Civita connection, which is uniquely defined by the following two properties:

$$ \nabla^g g = 0, \quad \text{and} \quad \mathbf{T}^g = 0, $$

where $\mathbf{T}^g := \mathbf{T}^{\nabla^g}$ is the torsion of $\nabla^g$. The Levi-Civita connection is determined by the equations

$$ \Gamma^{\kappa}_{\lambda\mu} = \frac{g^{\kappa\rho}}{2} (g_{\rho\lambda,\mu} + g_{\rho\mu,\lambda} - g_{\lambda\mu,\rho}), $$

that can be inverted as follows:

$$ g_{\kappa\lambda,\mu} = g_{\kappa\rho} \Gamma^{\rho}_{\lambda\mu} + g_{\lambda\sigma} \Gamma^{\sigma}_{\kappa\mu}. $$

One concludes therefore that, $g$ being invertible, the data of $\{g_{\mu\nu}, \partial_{\mu}g_{\kappa\lambda}, \partial_{\mu}\partial_{\nu}g_{\kappa\lambda}\}$ is equivalent to that of $\{g_{\mu\nu}, \Gamma^{\kappa}_{\lambda\mu}, \partial_{\nu}\Gamma^{\kappa}_{\lambda\mu}\}$.

### A.2. Action of the diffeomorphism group.
Following Erlangen's Program of Felix Klein [38], a geometry is defined by a group action. As such, Differential Geometry is defined by the action of the diffeomorphism group of a manifold. One of the main questions dealt with in this work is covariance under the diffeomorphism group. Let us describe how tensors and connections transform under the action of this group.

Let $\varphi: U \to V$ be a smooth map. The differential $T\varphi$ of $\varphi$ allows to push-forward a contravariant vector $\mathbf{X}$ at $m \in U$ to a contravariant vector $T_m\varphi \cdot \mathbf{X}$ at $\varphi(m)$. Globalizing this construction for a vector field $\mathbf{X}$ on $U$, one obtains a section $T\varphi \cdot \mathbf{X}$ of the pullback bundle

$$ \varphi^*TV := \coprod_{m \in U} T_{\varphi(m)}V. $$

When $\varphi$ is a diffeomorphism, for a vector field $\mathbf{X}$ on $U$, one can define a vector field $\varphi_*\mathbf{X}$ on $V$ by setting at a point $p \in V$:

$$ (\varphi_*\mathbf{X})_p := T_{\varphi^{-1}(p)}\varphi \cdot \mathbf{X}_{\varphi^{-1}(p)}. $$

One can extends this construction diagonally to higher order contravariant tensors in order to obtain a linear map $\varphi_*$ from contravariant tensor fields on $U$ to contravariant tensor fields on $V$.

Let now $\omega$ be a covariant tensor field on $V$ (or a scalar field). For a smooth map $\varphi: U \to V$, there is a natural covariant tensor field of the same type on $U$, obtained by pullback by $\varphi$. It is given at a point $m \in U$ by:

$$ (\varphi^*\alpha)_m(\mathbf{X}_1, \dots, \mathbf{X}_k) := \alpha_{\varphi(m)}(T_m\varphi \cdot \mathbf{X}_1, \dots, T_m\varphi \cdot \mathbf{X}_k). $$

For 1-forms, using $\star$ for the dual transpose, this action can also be written as

$$ \varphi^*\alpha = (T\varphi)^* \cdot \alpha \circ \varphi. $$

For exact 1-forms, one gets

$$ (\varphi^* \mathrm{d} f)_m(\mathbf{X}) = \mathrm{d} f_{\varphi(m)}(T_m\varphi \cdot \mathbf{X}) = \mathrm{d}(f \circ \varphi)_m \cdot \mathbf{X} = \mathrm{d}(\varphi^* f)_m \cdot \mathbf{X}, $$

which we sum up as $\varphi^* \mathrm{d} = \mathrm{d} \varphi^*$ (exterior derivatives and pullbacks commute).

When $\varphi$ is a diffeomorphism, one can also define the pullback to $U$ of a contravariant tensor field $\mathbf{X}$ on $V$ as the pushforward by the diffeomorphism $\varphi^{-1}$. For instance for a vector field:

$$ (\varphi^* \mathbf{X})_m := ((\varphi^{-1})_* \mathbf{X})_m = T_{\varphi(m)} \varphi^{-1} \cdot \mathbf{X}_{\varphi(m)}. $$
In the same line of thought, one can define the pushforward of a covariant tensor field $\omega$, by pulling back by the inverse diffeomorphism as in $\varphi_*\omega = (\varphi^{-1})^*\omega$.

The right action $\varphi^*$ of the diffeomorphism group (the pullback) and its inverse left action $\varphi_*$ (the pushforward) can be extended diagonally to arbitrary mixed tensor fields by using the tensor product decomposition of the tensor bundle. As intended, the pullback and the pushforward commute with contractions.

**Example A.3.** The pullback of a metric $g: \mathcal{X}^1(\mathcal{M}) \to \Omega^1(\mathcal{M})$ is given by

$$ (\varphi^*g)_m = (T\varphi)^* \cdot g_{\varphi(m)} \cdot T\varphi, $$

and the pullback of a cometric $g^{-1}: \Omega^1(\mathcal{M}) \to \mathcal{X}^1(\mathcal{M})$ is given by

$$ (\varphi^*g^{-1})_m := (T\varphi)^{-1} \cdot (g^{-1})_{\varphi(m)} \cdot (T\varphi)^{-*}. $$

One infers for instance that the gradient vector field $\mathrm{grad}^g f := g^{-1} \mathrm{d}f$ is general covariant of the couple of variables $(g, f)$. This means that:

$$
\begin{align*}
\mathrm{grad}^{\varphi^*g} \varphi^* f &= (\varphi^*g)^{-1} \mathrm{d}(\varphi^* f) \\
&= T\varphi^{-1} \cdot g^{-1} \cdot T\varphi^{-*} \cdot T\varphi^* \cdot \mathrm{d}f \\
&= (T\varphi)^{-1} \cdot g^{-1} \mathrm{d}f \\
&= \varphi^*(g^{-1} \mathrm{d}f) \\
&= \varphi^*(\mathrm{grad}^g f).
\end{align*}
$$

As already mentioned, covariant derivatives are not tensor fields. However, it is also possible to define pullbacks of covariant derivatives. This definition, in order to be natural, is required to satisfy

$$ \varphi^*(\nabla_X Y) = (\varphi^*\nabla)_{\varphi^*X} (\varphi^*Y). $$

which yields to the following definition of the pullback of a covariant derivative:

$$ (A.1) \qquad (\varphi^*\nabla)_X Y := \varphi^* (\nabla_{\varphi_*X} \varphi_* Y). $$

**Lemma A.4.** The pullback of a covariant derivative is still a covariant derivative. Moreover, for any tensor field $t$, one has

$$ \varphi^*(\nabla_X t) = (\varphi^*\nabla)_{\varphi^*X} (\varphi^*t). $$

*Proof.* Recall that a covariant derivative coincides with $\mathrm{d}$ on scalar fields, and that once it is defined on contravariant vector fields, it extends in a unique way to the tensor bundle using Leibniz rule. Let us thus check that (A.1) defines indeed a covariant derivative on $T\mathcal{M}$. One has

$$ (\varphi^*\nabla)_{f\mathbf{X}}(\mathbf{Y}) = \varphi^*(\nabla_{\varphi_*(f\mathbf{X})}(\varphi_*\mathbf{Y})) = \varphi^*(\varphi_*f)\varphi^*(\nabla_{\varphi_*\mathbf{X}}(\varphi_*\mathbf{Y})) = f(\varphi^*\nabla)_{\mathbf{X}}(\mathbf{Y}), $$

and

$$ (\varphi^*\nabla)_X(f\mathbf{Y}) = \varphi^*(\nabla_{\varphi_*X}\varphi_*(f\mathbf{Y})) = \varphi^*(\nabla_{\varphi_*X}(\varphi_*f\varphi_*\mathbf{Y})) = \varphi^*(\mathrm{d}\varphi_*f(\mathbf{X})\varphi_*\mathbf{Y} + \varphi_*f\nabla_{\varphi_*X}(\varphi_*\mathbf{Y})). $$

Recall that $\varphi^* \mathrm{d} = \mathrm{d}\varphi^*$, hence:

$$ (\varphi^*\nabla)_X(f\mathbf{Y}) = \mathrm{d}f(\mathbf{X})\mathbf{Y} + f(\varphi^*\nabla)_X(\mathbf{Y}). $$

Similarly the action of diffeomorphisms is the composition on scalar fields, and once it is defined on contravariant vector fields, it extends in a unique way to the tensor bundle using the diagonal action on contravariant tensor fields and the compatibility with contraction. We have already checked that for scalar fields $\varphi^* \mathrm{d} = \mathrm{d}\varphi^*$ and (A.1) is an ad-hoc definition to ensure that $\varphi^*\nabla = (\varphi^*\nabla)\varphi^*$ on contravariant vector fields. The extension of this formula to all tensor fields is then formal (see Example A.5 below). $\square$
**Example A.5.** For a metric $g$, one has:

$$ 
\begin{aligned}
(\varphi^*\nabla)(\varphi^*g)(\varphi^*\mathbf{X}, \varphi^*\mathbf{Y})
&= d\varphi^* (g(\mathbf{X}, \mathbf{Y})) - (\varphi^*g)((\varphi^*\nabla)\varphi^*\mathbf{X}, \varphi^*\mathbf{Y}) - (\varphi^*g)(\varphi^*\mathbf{X}, (\varphi^*\nabla)\varphi^*\mathbf{Y}) \\
&= \varphi^* d(g(\mathbf{X}, \mathbf{Y})) - (\varphi^*g)(\varphi^*(\nabla\mathbf{X}), \varphi^*\mathbf{Y}) - (\varphi^*g)(\varphi^*\mathbf{X}, \varphi^*(\nabla\mathbf{Y})) \\
&= \varphi^* (d(g(\mathbf{X}, \mathbf{Y}))) - g(\nabla\mathbf{X}, \mathbf{Y}) - g(\mathbf{X}, \nabla\mathbf{Y})) \\
&= \varphi^* ((\nabla g)(\mathbf{X}, \mathbf{Y})) = \varphi^*(\nabla g)(\varphi^*\mathbf{X}, \varphi^*\mathbf{Y}).
\end{aligned}
$$

We list below some formal consequences of (A.1), without demonstration.

**Corollary A.6.** Let $\nabla$ be a covariant derivative. The following properties hold, for any local diffeomorphism $\varphi$ of $\mathcal{M}$.

(1) General Covariance of the torsion tensor:

$$ \mathbf{T}^{\varphi^*\nabla} = \varphi^* (\mathbf{T}^{\nabla}). $$

(2) General Covariance of the curvature tensor:

$$ \mathbf{R}^{\varphi^*\nabla} = \varphi^* (\mathbf{R}^{\nabla}). $$

(3) General Covariance of the Hessian tensor:

$$ \mathbf{Hess}^{\varphi^*\nabla} \varphi^* f = \varphi^* (\mathbf{Hess}^{\nabla} f). $$

Recall that on a pseudo-Riemannian manifold with metric $g$, we denote the Levi-Civita connection by $\nabla^g$. The connection $\varphi^*\nabla^g$ satisfies $\mathbf{T}^{\varphi^*\nabla^g} = \varphi^* (\mathbf{T}^{\nabla^g}) = 0$, and $(\varphi^*\nabla^g)\varphi^*g = \varphi^*(\nabla^g g) = 0$. Therefore, by the unique characterization of the Levi-Civita connection, we have

$$ \nabla^{\varphi^*g} = \varphi^*\nabla^g. $$

We can hence specialize the above results as follow:

$$ \mathbf{R}^{\varphi^*g} = \varphi^*\mathbf{R}^g, \quad \text{and} \quad \mathbf{Hess}^{\varphi^*g} \varphi^* f = \varphi^* (\mathbf{Hess}^g f). $$

It follows that the Ricci tensor $\mathbf{Ric}_g := (R_{\lambda\nu})$ where $R_{\lambda\nu} := R^\kappa_{\lambda\kappa\nu}$ and the scalar curvature $R_g := g^{\lambda\nu}R_{\lambda\nu}$, obtained by contraction, also satisfy

$$ \mathbf{Ric}_{\varphi^*g} = \varphi^*\mathbf{Ric}_g, \quad \text{and} \quad R_{\varphi^*g} = \varphi^*R_g. $$

**A.3. Normalization of the Christoffel symbols.** We consider the Christoffel symbols of a connection $\nabla$ in local coordinates

$$ \Gamma^\kappa_{\lambda\mu} = d x^\kappa (\nabla_{\partial_\mu} \partial_\lambda). $$

and their jets. In this section, for a local diffeomorphism $\varphi$, we adopt the notation

$$ \bar{\Gamma}^\kappa_{\lambda\mu} := d x^\kappa ((\varphi^*\nabla)_{\partial_\mu} \partial_\lambda), $$

and we shall establish the following enhancement of a classical normalization result (see also [61, (35.8)] and [37, Section 2]).

**Theorem A.7 (Normalization of the Christoffel symbols).** For any connection $\nabla$, and for any point $m \in \mathcal{M}$, there exists a diffeomorphism with trivial linear part such that, punctually at $m$,

$$ \bar{\Gamma}^\kappa_{(\lambda\mu)}(m) = 0 \quad \text{and} \quad \bar{\Gamma}^\kappa_{(\lambda\mu,\nu)}(m) = 0, $$

where the comma in the index is the traditional notation for the partial derivative, and where parentheses is a shorthand notation for symmetrization.

*Proof.* Let us first set the context of the proof. We consider a local diffeomorphism $\varphi$ fixing a local chart around $m$, equipped with a local coordinate system $(x^\mu)$. At $\varphi(m)$, we define the Taylor coefficients

$$ A^\kappa_\lambda = \varphi_*(\partial_\lambda \varphi^\kappa), \quad B^\kappa_{\lambda\mu} = \varphi_*(\partial_\mu \partial_\lambda \varphi^\kappa), \quad C^\kappa_{\lambda\mu\nu} = \varphi_*(\partial_\nu \partial_\mu \partial_\lambda \varphi^\kappa). $$
Note that the matrix $(A^\kappa_\lambda)$ is by assumption invertible, and that by virtue of the Schwarz Lemma, both $B^\kappa_{\lambda\mu}$ and $C^\kappa_{\lambda\mu\nu}$ are totally symmetric with respect to their lower indices. It follows from the very definitions of **A**, **B**, **C** that

$$ \left\{ \begin{array}{l} \nabla_{\varphi_* \partial_\mu} A^\kappa_\lambda = \nabla_{\varphi_* \partial_\mu} \varphi_*(\partial_\lambda \varphi^\kappa) = \varphi_*(\nabla_{\partial_\mu}(\partial_\lambda \varphi^\kappa)) = \varphi_*(\partial_\mu \partial_\lambda \varphi^\kappa) = B^\kappa_{\lambda\mu} \\ \nabla_{\varphi_* \partial_\nu} B^\kappa_{\lambda\mu} = \nabla_{\varphi_* \partial_\nu} \varphi_*(\partial_\mu \partial_\lambda \varphi^\kappa) = \varphi_*(\nabla_{\partial_\nu}(\partial_\mu \partial_\lambda \varphi^\kappa)) = \varphi_*(\partial_\nu \partial_\mu \partial_\lambda \varphi^\kappa) = C^\kappa_{\lambda\mu\nu} \end{array} \right. . $$

We can now compute $\bar{\Gamma}$. By definition:

$$ (A.2) \qquad (\varphi_*(\partial_\lambda)\varphi(m)) = (\partial_\lambda\varphi^\kappa)_m(\partial_\kappa)\varphi(m) = (A^\kappa_\lambda\partial_\kappa)\varphi(m). $$

By Leibniz rule, at a point $\varphi(m)$, one has

$$ \begin{aligned} \nabla_{\varphi_* \partial_\mu} (\varphi_* \partial_\lambda) &= \nabla_{\varphi_* \partial_\mu} (A^\sigma_\lambda \partial_\sigma) = \nabla_{\varphi_* \partial_\mu} (\partial_\sigma) A^\sigma_\lambda + \nabla_{\varphi_* \partial_\mu} (A^\sigma_\lambda) \partial_\sigma \\ &= \nabla_{\partial_\tau} (\partial_\sigma) A^\tau_\mu A^\sigma_\lambda + B^\sigma_{\lambda\mu} \partial_\sigma = (\Gamma^\rho_{\sigma\tau} A^\tau_\mu A^\sigma_\lambda + B^\rho_{\lambda\mu}) \partial_\rho. \end{aligned} $$

Therefore, inverting (A.2) to get

$$ \partial_\rho = (\mathbf{A}^{-1})^\kappa_\rho \varphi_* \partial_\kappa, $$

and pulling back at $m$:

$$ (\varphi^* \nabla) \partial_\mu (\partial_\lambda) = \varphi^* (\nabla_{\varphi_*} \partial_\mu (\varphi_* \partial_\lambda)) = \varphi^* \left( (\mathbf{A}^{-1})^\kappa_\rho \left( \Gamma^\rho_{\sigma\tau} A^\tau_\mu A^\sigma_\lambda + B^\rho_{\lambda\mu} \right) \right) \partial_\kappa. $$

We get the following expression for $\bar{\Gamma}$,

$$ (A.3) \qquad \bar{\Gamma}^\kappa_{\lambda\mu}(m) = (\mathbf{A}^{-1})^\kappa_\rho \left( \Gamma^\rho_{\sigma\tau} A^\tau_\mu A^\sigma_\lambda + B^\rho_{\lambda\mu} \right) (\varphi(m)). $$

Differentiating this functional equality, one gets first

$$ \bar{\Gamma}^\kappa_{\lambda\mu,\nu}(m) = (\varphi_* \partial_\nu) \left( (\mathbf{A}^{-1})^\kappa_\rho \left( \Gamma^\rho_{\sigma\tau} A^\tau_\mu A^\sigma_\lambda + B^\rho_{\lambda\mu} \right) \right) (\varphi(m)), $$

from which one infers, using Leibniz rule and the definitions of **A**, **B**, **C**:

$$ \begin{aligned} \bar{\Gamma}^\kappa_{\lambda\mu,\nu}(m) = (\mathbf{A}^{-1})^\kappa_\alpha B^{\alpha}_{\beta\nu}(\mathbf{A}^{-1})^\beta_\rho & \left( \Gamma^\rho_{\sigma\tau} A^\tau_\mu A^\sigma_\lambda + B^\rho_{\lambda\mu} \right) (\varphi(m)) \\ & + (\mathbf{A}^{-1})^\kappa_\rho \left( \Gamma^\rho_{\sigma\tau,\nu} A^\nu_\mu A^\sigma_\lambda + \Gamma^\rho_{\sigma\tau} A^\tau_\mu B^\sigma_{\lambda\nu} + \Gamma^\rho_{\sigma\tau} B^\tau_{\mu\nu} A^\sigma_\lambda + C^\rho_{\lambda\mu\nu} \right) (\varphi(m)). \end{aligned} $$

We sum up the results obtained so far in the special case where $\varphi(m) = m$ and $T\varphi = \text{id}$. At the point $m$, one has:

$$ (A.4) \qquad \left\{ \begin{array}{l} \bar{\Gamma}^\kappa_{\lambda\mu} = \Gamma^\kappa_{\lambda\mu} + B^\kappa_{\lambda\mu} \\ \bar{\Gamma}^\kappa_{\lambda\mu,\nu} = \Gamma^\kappa_{\lambda\mu,\nu} + B^\kappa_{\rho\nu} (\Gamma^\rho_{\lambda\mu} + B^\rho_{\lambda\mu}) + \Gamma^\kappa_{\sigma\mu} B^\sigma_{\lambda\nu} + \Gamma^\kappa_{\lambda\tau} B^\tau_{\mu\nu} + C^\kappa_{\lambda\mu\nu} \end{array} \right. . $$

Since **B** and **C** are arbitrary symmetric in lower indices, we can fix them one-by-one in order to cancel out the symmetric parts (firstly **B** in the top row and then **C** in the bottom row). $\square$

In the case of the Levi-Civita connection, the normalization result becomes

**Corollary A.8.** For any point $m \in (\mathcal{M}, g)$, there exists a diffeomorphism with trivial linear part $\varphi$ such that, punctually at $m$, the Levi-Civita connection of the metric $\varphi^*g$ satisfies

$$ \Gamma^\kappa_{\lambda\mu} = 0 \quad \text{and} \quad \Gamma^\kappa_{\lambda\mu,\nu} = -\frac{1}{3} (R^\kappa_{\lambda\mu\nu} + R^\kappa_{\mu\lambda\nu}). $$

*Proof.* The Levi-Civita connection is symmetric, hence

$$ \bar{\Gamma}^\kappa_{(\lambda\mu)} = 0 \iff \bar{\Gamma}^\kappa_{\lambda\mu} = 0. $$

Moreover, at a point where the Christoffel coefficients vanish, the local components of the Riemann tensor are $R^\kappa_{\lambda\mu\nu} = \Gamma^\kappa_{\nu\lambda,\mu} - \Gamma^\kappa_{\mu\lambda,\nu}$. One infers, using the symmetries of $\Gamma^\kappa_{\mu\lambda,\nu}$ in $\mu, \lambda$ that

$$ R^\kappa_{\lambda\mu\nu} + R^\kappa_{\mu\lambda\nu} = \Gamma^\kappa_{\nu\lambda,\mu} + \Gamma^\kappa_{\nu\mu,\lambda} - 2\Gamma^\kappa_{\mu\lambda,\nu} = \Gamma^\kappa_{(\mu\lambda,\nu)} - 3\Gamma^\kappa_{\mu\lambda,\nu}. $$
We finally get that

$$ \Gamma^{\kappa}_{(\lambda\mu,\nu)} = 0 \iff \Gamma^{\kappa}_{\lambda\mu,\nu} = -\frac{1}{3}(R^{\kappa}_{\lambda\mu\nu} + R^{\kappa}_{\mu\lambda\nu}), $$

which ends the proof.

<figure><img src="image_5.png" /></figure>

## APPENDIX B. VOLUME FORMS AND HODGE OPERATOR

### B.1. Volume pseudo-form.
The manifold $\mathcal{M}$ is orientable if and only if there exists a global trivialization $\vartheta$ of the line bundle $\Lambda^4 T\mathcal{M}$. Assume now that $\mathcal{M}$ is oriented. Given a Lorentzian metric $g$ on $\mathcal{M}$, one can then define the Riemannian volume form $\text{vol}_g$ as follows (the dependence in $\vartheta$ is implicit). At a point $m \in \mathcal{M}$, given the orientation $\vartheta_m$ of $T_m\mathcal{M}$ (seen as a non-zero element of the 1-dimensional vector space $\Lambda^4 T_m\mathcal{M}$), a basis $v := (v_0, v_1, v_2, v_3)$ of $T_m\mathcal{M}$ is called direct if

$$ (v^0 \wedge v^1 \wedge v^2 \wedge v^3)(\vartheta_m) > 0, $$

where $(v^0, v^1, v^2, v^3)$ is the dual basis of $T_m\mathcal{M}^*$. Given a direct orthonormal basis $(v_0, v_1, v_2, v_3)$ of $T_m\mathcal{M}$, we set

$$ \text{vol}_g(m) := v^0 \wedge v^1 \wedge v^2 \wedge v^3. $$

Since any other direct orthonormal basis is written $Lv$, where $L$ belongs to the special Lorentz group

$$ \text{SO}(3, 1) := \{L \in \text{O}(3, 1); \det L = 1\}, $$

this construction does not depend on the choice of the direct orthonormal basis $v = (v_0, v_1, v_2, v_3)$. The data $\text{vol}_g(m)$ are glued together in order to define a global smooth volume form $\text{vol}_g \in \Omega^4(\mathcal{M})$.

For a local diffeomorphism $\varphi: U \to \bar{U}$, recall from subsection A.2 that

$$ \varphi^*g(\varphi^*v_i, \varphi^*v_j) = g(v_i, v_j) \quad \text{and} \quad \varphi^*\vartheta(\varphi^*v) = \vartheta(v). $$

Therefore, if $v$ is a direct orthonormal basis of $T_{\varphi(m)}\mathcal{M}$ for $(g, \vartheta)$, then $\varphi^*v$ is a direct orthonormal basis of $T_m\mathcal{M}$ for $(\varphi^*g, \varphi^*\vartheta)$. Evaluating the volume forms on $\varphi^*v$, one infers that

$$ \text{vol}_{\varphi^*g, \varphi^*\vartheta} = \det(T\varphi)^{-1}(\text{vol}_{g,\vartheta}) \circ \varphi^{-1} = \varphi^* \text{vol}_{g,\vartheta}. $$

In other words the volume form $\text{vol}_{g,\vartheta}$ is a covariant of the pair $(g, \vartheta)$ under the group of all diffeomorphisms.

If the orientation $\vartheta$ has been fixed, a slight modification of our arguments shows that $\text{vol}_g$ is a covariant of the field $g$ under the group of orientation-preserving diffeomorphisms, but that it changes sign when the orientation is reversed (it is a density):

$$ \text{vol}_{\varphi^*g, \vartheta} = |\det(T\varphi)|^{-1}(\text{vol}_{g,\vartheta}) \circ \varphi^{-1}. $$

In the present work we consider only the subgroup of orientation-preserving diffeomorphisms of $\mathcal{M}$, which is common in General Relativity.

### B.2. Hodge star operator.
We consider an oriented Lorentzian 4-dimensional manifold $(\mathcal{M}, g, \vartheta)$ with metric signature $(- + + +)$. We fix the orientation $\vartheta$ and from now on we will only imply it. Denote by $*_g$ the Hodge star operator. Recall that it is $\mathcal{C}^\infty(\mathcal{M})$-linear and defined recursively on decomposable $k$ forms by the normalization $*_g 1 = \text{vol}_g$ and by the formula

$$ i_V(*_g \omega) = *_g(\omega \wedge V^b), $$

where $i_V$ is the contraction by a vector field $V$. Detailing the definition of $*_g$, one has for every 1-forms $\omega^1, \omega^2, \omega^3, \omega^4$:

$$ \begin{align} \text{vol}_g(\omega^1\sharp, \omega^2\sharp, \omega^3\sharp, \omega^4\sharp) &= *_g(1)(\omega^1\sharp, \omega^2\sharp, \omega^3\sharp, \omega^4\sharp) = *_g(\omega^1)(\omega^2\sharp, \omega^3\sharp, \omega^4\sharp) \nonumber \\ &= *_g(\omega^1 \wedge \omega^2)(\omega^3\sharp, \omega^4\sharp) \nonumber \\ &= *_g(\omega^1 \wedge \omega^2 \wedge \omega^3)(\omega^4\sharp) \nonumber \\ &= *_g(\omega^1 \wedge \omega^2 \wedge \omega^3 \wedge \omega^4). \tag{B.1} \end{align} $$
*Remark B.3.* Whereas the vector fields $d\Psi^I$ are naturally covariant under the group of general diffeomorphisms, the supplementary covector field $*_g(d\Psi^1 \wedge d\Psi^2 \wedge d\Psi^3)$ changes sign when the orientation is reversed. It could be made covariant under the group of general diffeomorphisms by making its dependence on the orientation explicit—it would become a tensor density, see *subsection B.1* for more details on the volume pseudo-form—, but we have the feeling that this would hinder the clarity of the exposition.

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(Mina Chapon) **UNIVERSITÉ PARIS-SACLAY, ENS PARIS-SACLAY, CENTRALESUPÉLEC, CNRS, LMPS -**
**LABORATOIRE DE MÉCANIQUE PARIS-SACLAY, 91190, GIF-SUR-YVETTE, FRANCE**

Email address: mina.chapon@ens-paris-saclay.fr

(Lionel Darondeau) **SORBONNE UNIVERSITÉ, CNRS, IMJ-PRG, F-75005 PARIS, FRANCE.**

Email address: darondeau@imj-prg.fr

(Rodrigue Desmorat) **UNIVERSITÉ PARIS-SACLAY, ENS PARIS-SACLAY, CENTRALESUPÉLEC, CNRS, LMPS -**
**LABORATOIRE DE MÉCANIQUE PARIS-SACLAY, 91190, GIF-SUR-YVETTE, FRANCE**

Email address: rodrigue.desmorat@ens-paris-saclay.fr

(Clément Ecker) **UNIVERSITÉ PARIS-SACLAY, ENS PARIS-SACLAY, CENTRALESUPÉLEC, CNRS, LMPS -**
**LABORATOIRE DE MÉCANIQUE PARIS-SACLAY, 91190, GIF-SUR-YVETTE, FRANCE**

Email address: clement.ecker@ens-paris-saclay.fr

(Boris Kolev) **UNIVERSITÉ PARIS-SACLAY, ENS PARIS-SACLAY, CENTRALESUPÉLEC, CNRS, LMPS - LABORATOIRE DE MÉCANIQUE PARIS-SACLAY, 91190, GIF-SUR-YVETTE, FRANCE**

Email address: boris.kolev@ens-paris-saclay.fr