\documentclass[aps,prl,reprint,longbibliography,floatfix]{revtex4-1} \usepackage[utf8]{inputenc} \usepackage{amsmath,graphicx,upgreek,amssymb,xcolor} \usepackage[colorlinks=true,urlcolor=purple,citecolor=purple,filecolor=purple,linkcolor=purple]{hyperref} \usepackage[english]{babel} \makeatletter % A change to a babel macro -- Don't ask! \def\bbl@set@language#1{% \edef\languagename{% \ifnum\escapechar=\expandafter`\string#1\@empty \else\string#1\@empty\fi}% %%%% ADDITION \@ifundefined{babel@language@alias@\languagename}{}{% \edef\languagename{\@nameuse{babel@language@alias@\languagename}}% }% %%%% END ADDITION \select@language{\languagename}% \expandafter\ifx\csname date\languagename\endcsname\relax\else \if@filesw \protected@write\@auxout{}{\string\select@language{\languagename}}% \bbl@for\bbl@tempa\BabelContentsFiles{% \addtocontents{\bbl@tempa}{\xstring\select@language{\languagename}}}% \bbl@usehooks{write}{}% \fi \fi} % The user interface \newcommand{\DeclareLanguageAlias}[2]{% \global\@namedef{babel@language@alias@#1}{#2}% } \makeatother \DeclareLanguageAlias{en}{english} \newcommand{\brad}[1]{{\color{red} #1}} % Our mysterious boy \def\urusi{URu$_{\text2}$Si$_{\text2}$} \def\e{{\text{\textsc{elastic}}}} % "elastic" \def\i{{\text{\textsc{int}}}} % "interaction" \def\Dfh{D$_{\text{4h}}$} % Irreducible representations (use in math mode) \def\Aog{{\text A_{\text{1g}}}} \def\Atg{{\text A_{\text{2g}}}} \def\Bog{{\text B_{\text{1g}}}} \def\Btg{{\text B_{\text{2g}}}} \def\Eg {{\text E_{\text g}}} \def\Aou{{\text A_{\text{1u}}}} \def\Atu{{\text A_{\text{2u}}}} \def\Bou{{\text B_{\text{1u}}}} \def\Btu{{\text B_{\text{2u}}}} \def\Eu {{\text E_{\text u}}} % Variables to represent some representation \def\X{\text X} \def\Y{\text Y} % Units \def\J{\text J} \def\m{\text m} \def\K{\text K} \def\GPa{\text{GPa}} \def\A{\text{\r A}} % Other \def\op{\textsc{op}} % order parameter \def\ho{\textsc{ho}} % hidden order \def\rus{\textsc{rus}} % resonant ultrasound spectroscopy \def\Rus{\textsc{Rus}} % Resonant ultrasound spectroscopy \def\afm{\textsc{afm}} % antiferromagnetism \def\recip{{\{-1\}}} % functional reciprocal \begin{document} \title{Elastic properties of hidden order in \urusi\ are reproduced by a staggered nematic} \author{Jaron Kent-Dobias} \author{Michael Matty} \author{Brad Ramshaw} \affiliation{ Laboratory of Atomic \& Solid State Physics, Cornell University, Ithaca, NY, USA } \date\today \begin{abstract} We develop a phenomenological mean field theory describing the hidden order phase in \urusi\ as a nematic of the $\Bog$ representation staggered along the $c$-axis. Several experimental features are reproduced by this theory: the topology of the temperature--pressure phase diagram, the response of the elastic modulus $(C_{11}-C_{12})/2$ above the transition at ambient pressure, and orthorhombic symmetry breaking in the high-pressure antiferromagnetic phase. In this scenario, hidden order is characterized by broken rotational symmetry that is modulated along the $c$-axis, the primary order of the high-pressure phase is an unmodulated nematic, and the triple point joining those two phases with the high-temperature paramagnetic phase is a Lifshitz point. \end{abstract} \maketitle \emph{Introduction.} \urusi\ is a paradigmatic example of a material with an ordered state whose broken symmetry remains unknown. This state, known as \emph{hidden order} (\ho), sets the stage for unconventional superconductivity that emerges at even lower temperatures. At sufficiently large hydrostatic pressures, both superconductivity and \ho\ give way to local moment antiferromagnetism (\afm) \cite{hassinger_temperature-pressure_2008}. Despite over thirty years of effort, the symmetry of the \ho\ state remains unknown, and modern theories \cite{kambe_odd-parity_2018, haule_arrested_2009, kusunose_hidden_2011, kung_chirality_2015, cricchio_itinerant_2009, ohkawa_quadrupole_1999, santini_crystal_1994, kiss_group_2005, harima_why_2010, thalmeier_signatures_2011, tonegawa_cyclotron_2012, rau_hidden_2012, riggs_evidence_2015, hoshino_resolution_2013, ikeda_theory_1998, chandra_hastatic_2013, harrison_hidden_nodate, ikeda_emergent_2012} propose a variety of possibilities. This work seeks to unify two experimental observations: first, the $\Bog$ ``nematic" elastic susceptibility $(C_{11}-C_{12})/2$ softens anomalously from room temperature down to $T_{\text{\ho}}=17.5~$ K \cite{de_visser_thermal_1986}; and second, a $\Bog$ nematic distortion is observed by x-ray scattering under sufficient pressure to destroy the \ho\ state \cite{choi_pressure-induced_2018}. Recent resonant ultrasound spectroscopy (\rus) measurements were used to examine the thermodynamic discontinuities in the elastic moduli at $T_{\text{\ho}}$ \cite{ghosh_single-component_nodate}. The observation of discontinues only in compressional, or $\Aog$, elastic moduli requires that the point-group representation of \ho\ be one-dimensional. This rules out many order parameter candidates \cite{thalmeier_signatures_2011, tonegawa_cyclotron_2012, rau_hidden_2012, riggs_evidence_2015, hoshino_resolution_2013, ikeda_emergent_2012, chandra_origin_2013} in a model-independent way, but doesn't differentiate between those that remain. Recent x-ray experiments discovered rotational symmetry breaking in \urusi\ under pressure \cite{choi_pressure-induced_2018}. Above 0.13--0.5 $\GPa$ (depending on temperature), \urusi\ undergoes a $\Bog$ nematic distortion. While it remains unclear as to whether this is a true thermodynamic phase transition, it may be related to the anomalous softening of the $\Bog$ elastic modulus $(C_{11}-C_{12})/2$ that occurs over a broad temperature range at zero-pressure \cite{wolf_elastic_1994, kuwahara_lattice_1997}. Motivated by these results---which hint at a $\Bog$ strain susceptibility associated with the \ho\ state---we construct a phenomenological mean field theory for an arbitrary \op\ coupled to strain, and the determine the effect of its phase transitions on the elastic response in different symmetry channels. We find that only one \op\ symmetry reproduces the anomalous $\Bog$ elastic modulus, which softens in a Curie--Weiss-like manner from room temperature but cusps at $T_{\text{\ho}}$. That theory associates \ho\ with a $\Bog$ \op\ modulated along the $c$-axis, the \afm\ state with uniform $\Bog$ order, and the triple point between them with a Lifshitz point. Besides the agreement with ultrasound data across a broad temperature range, the theory predicts uniform $\Bog$ strain at high pressure---the same distortion that was recently seen in x-ray scattering experiments \cite{choi_pressure-induced_2018}. This theory strongly motivates future ultrasound experiments under pressure approaching the Lifshitz point, which should find that the $(C_{11}-C_{12})/2$ modulus diverges as the uniform $\Bog$ strain of the \afm\ phase is approached. \emph{Model.} The point group of \urusi\ is \Dfh, and any coarse-grained theory must locally respect this symmetry in the high-temperature phase. Our phenomenological free energy density contains three parts: the elastic free energy, the \op, and the interaction between strain and \op. The most general quadratic free energy of the strain $\epsilon$ is $f_\e=C^0_{ijkl}\epsilon_{ij}\epsilon_{kl}$ \footnote{Components of the elastic modulus tensor $C$ were given in the popular Voigt notation in the abstract and introduction. Here and henceforth the notation used is that natural for a rank-four tensor.}. Linear combinations of the six independent components of strain form five irreducible components of strain as \begin{equation} \begin{aligned} & \epsilon_{\Aog,1}=\epsilon_{11}+\epsilon_{22} \hspace{0.15\columnwidth} && \epsilon_\Bog=\epsilon_{11}-\epsilon_{22} \\ & \epsilon_{\Aog,2}=\epsilon_{33} && \epsilon_\Btg=2\epsilon_{12} \\ & \epsilon_\Eg=2\{\epsilon_{11},\epsilon_{22}\}. \end{aligned} \label{eq:strain-components} \end{equation} All quadratic combinations of these irreducible strains that transform like $\Aog$ are included in the free energy, \begin{equation} f_\e=\frac12\sum_\X C^0_{\X,ij}\epsilon_{\X,i}\epsilon_{\X,j}, \end{equation} where the sum is over irreducible representations of the point group and the bare elastic moduli $C^0_\X$ are \begin{equation} \begin{aligned} & C^0_{\Aog,11}=\tfrac12(C^0_{1111}+C^0_{1122}) && C^0_{\Bog}=\tfrac12(C^0_{1111}-C^0_{1122}) \\ & C^0_{\Aog,22}=C^0_{3333} && C^0_{\Btg}=C^0_{1212} \\ & C^0_{\Aog,12}=C^0_{1133} && C^0_{\Eg}=C^0_{1313}. \end{aligned} \end{equation} The interaction between strain and an \op\ $\eta$ depends on the point group representation of $\eta$. If this representation is $\X$, the most general coupling to linear order is \begin{equation} f_\i=-b^{(i)}\epsilon_\X^{(i)}\eta. \end{equation} If there doesn't exist a component of strain that transforms like the representation $\X$ there can be no linear coupling, and the effect of the \op\ condensing at a continuous phase transition is to produce a jump in the $\Aog$ elastic moduli if $\eta$ is single-component \cite{luthi_sound_1970, ramshaw_avoided_2015, shekhter_bounding_2013}, and jumps in other elastic moduli if multicomponent \cite{ghosh_single-component_nodate}. Because we are interested in physics that anticipates the phase transition, we will focus our attention on \op s that can produce linear couplings to strain. Looking at the components present in \eqref{eq:strain-components}, this rules out all of the u-reps (which are odd under inversion) and the $\Atg$ irrep. If the \op\ transforms like $\Aog$ (e.g. a fluctuation in valence number), odd terms are allowed in its free energy and any transition will be first order and not continuous without fine-tuning. Since the \ho\ phase transition is second-order \cite{de_visser_thermal_1986}, we will henceforth rule out $\Aog$ \op s as well. For the \op\ representation $\X$ as any of those remaining---$\Bog$, $\Btg$, or $\Eg$---the most general quadratic free energy density is \begin{equation} \begin{aligned} f_\op=\frac12\big[&r\eta^2+c_\parallel(\nabla_\parallel\eta)^2 +c_\perp(\nabla_\perp\eta)^2 \\ &\qquad\qquad\qquad\quad+D_\perp(\nabla_\perp^2\eta)^2\big]+u\eta^4, \end{aligned} \label{eq:fo} \end{equation} where $\nabla_\parallel=\{\partial_1,\partial_2\}$ transforms like $\Eu$, and $\nabla_\perp=\partial_3$ transforms like $\Atu$. Other quartic terms are allowed---especially many for an $\Eg$ \op---but we have included only those terms necessary for stability when either $r$ or $c_\perp$ become negative. The full free energy functional of $\eta$ and $\epsilon$ is \begin{equation} \begin{aligned} F[\eta,\epsilon] &=F_\op[\eta]+F_\e[\epsilon]+F_\i[\eta,\epsilon] \\ &=\int dx\,(f_\op+f_\e+f_\i). \end{aligned} \label{eq:free_energy} \end{equation} Rather than analyze this two-argument functional directly, we begin by tracing out the strain and studying the behavior of \op\ alone. Later we will invert this procedure and trace out the \op\ when we compute the effective elastic moduli. The only strain relevant to the \op\ at linear coupling is $\epsilon_\X$, which can be traced out of the problem exactly in mean field theory. Extremizing the functional \eqref{eq:free_energy} with respect to $\epsilon_\X$ gives \begin{equation} 0 =\frac{\delta F[\eta,\epsilon]}{\delta\epsilon_\X(x)}\bigg|_{\epsilon=\epsilon_\star} =C^0_\X\epsilon^\star_\X(x)-b\eta(x), \end{equation} which in turn gives the strain field conditioned on the state of the \op\ field as $\epsilon_\X^\star[\eta](x)=(b/C^0_\X)\eta(x)$ at all spatial coordinates $x$, and $\epsilon_\Y^\star[\eta]=0$ for all other irreps $\Y\neq\X$. Upon substitution into the \eqref{eq:free_energy}, the resulting single-argument free energy functional $F[\eta,\epsilon_\star[\eta]]$ has a density identical to $f_\op$ with $r\to\tilde r=r-b^2/2C^0_\X$. \begin{figure}[htpb] \includegraphics[width=\columnwidth]{phase_diagram_experiments} \vspace{1em} \includegraphics[width=0.51\columnwidth]{phases_scalar}\hspace{-1.5em} \includegraphics[width=0.51\columnwidth]{phases_vector} \caption{ Phase diagrams for (a) \urusi\ from experiments (neglecting the superconducting phase) \cite{hassinger_temperature-pressure_2008} (b) mean field theory of a one-component ($\Bog$ or $\Btg$) Lifshitz point (c) mean field theory of a two-component ($\Eg$) Lifshitz point. Solid lines denote continuous transitions, while dashed lines denote first order transitions. Later, when we fit the elastic moduli predictions for a $\Bog$ \op\ to data along the ambient pressure line, we will take $\Delta\tilde r=\tilde r-\tilde r_c=a(T-T_c)$. } \label{fig:phases} \end{figure} With the strain traced out, \eqref{eq:fo} describes the theory of a Lifshitz point at $\tilde r=c_\perp=0$ \cite{lifshitz_theory_1942, lifshitz_theory_1942-1}. The properties discused below can all be found in a standard text, e.g.,~\cite{chaikin_principles_2000}. For a one-component \op\ ($\Bog$ or $\Btg$) and positive $c_\parallel$, it is traditional to make the field ansatz $\langle\eta(x)\rangle=\eta_*\cos(q_*x_3)$. For $\tilde r>0$ and $c_\perp>0$, or $\tilde r>c_\perp^2/4D_\perp$ and $c_\perp<0$, the only stable solution is $\eta_*=q_*=0$ and the system is unordered. For $\tilde r<0$ there are free energy minima for $q_*=0$ and $\eta_*^2=-\tilde r/4u$ and this system has uniform order with the \op\ symmetry, e.g., $\Bog$ or $\Btg$. For $c_\perp<0$ and $\tilde r0$, and the modulated phase is now characterized by helical order with $\langle\eta(x)\rangle=\eta_*\{\cos(q_*x_3),\sin(q_*x_3)\}$. The uniform to modulated transition is now continuous. This does not reproduce the physics of \ho, which has a first order transition between \ho\ and \afm, and so we will henceforth neglect the possibility of a multicomponent order parameter. The schematic phase diagrams for this model are shown in Figure~\ref{fig:phases}. \emph{Results.} We will now derive the effective elastic tensor $C$ that results from coupling of strain to the \op. The ultimate result, found in \eqref{eq:elastic.susceptibility}, is that $C_\X$ differs from its bare value $C^0_\X$ only for the symmetry $\X$ of the \op. Moreover, this modulus does not vanish at the unordered to modulated transition---as it would if the transition were a $q=0$ structural phase transition---but ends in a cusp. In this section we start by computing the susceptibility of the \op\ at the unordered to modulated transition, and then compute the elastic modulus for the same. The susceptibility of a single-component ($\Bog$ or $\Btg$) \op\ is \begin{equation} \begin{aligned} &\chi^\recip(x,x') =\frac{\delta^2F[\eta,\epsilon_\star[\eta]]}{\delta\eta(x)\delta\eta(x')}\bigg|_{\eta=\langle\eta\rangle} =\big[\tilde r-c_\parallel\nabla_\parallel^2 \\ &\qquad-c_\perp\nabla_\perp^2+D_\perp\nabla_\perp^4+12u\langle\eta(x)\rangle^2\big]\delta(x-x'), \end{aligned} \label{eq:sus_def} \end{equation} where $\recip$ indicates a functional reciprocal defined as \begin{equation} \int dx''\,\chi^\recip(x,x'')\chi(x'',x')=\delta(x-x'). \end{equation} Taking the Fourier transform and integrating out $q'$ gives \begin{equation} \chi(q) =\big(\tilde r+c_\parallel q_\parallel^2+c_\perp q_\perp^2+D_\perp q_\perp^4 +12u\sum_{q'}\langle\tilde\eta_{q'}\rangle\langle\tilde\eta_{-q'}\rangle\big)^{-1}. \end{equation} Near the unordered to modulated transition this yields \begin{equation} \begin{aligned} \chi(q) &=\big[c_\parallel q_\parallel^2+D_\perp(q_*^2-q_\perp^2)^2 +|\Delta\tilde r|\big]^{-1} \\ &=\frac1{D_\perp}\frac{\xi_\perp^4} {1+\xi_\parallel^2q_\parallel^2+\xi_\perp^4(q_*^2-q_\perp^2)^2}, \end{aligned} \label{eq:susceptibility} \end{equation} with $\xi_\perp=(|\Delta\tilde r|/D_\perp)^{-1/4}=\xi_{\perp0}|t|^{-1/4}$ and $\xi_\parallel=(|\Delta\tilde r|/c_\parallel)^{-1/2}=\xi_{\parallel0}|t|^{-1/2}$, where $t=(T-T_c)/T_c$ is the reduced temperature and $\xi_{\perp0}=(D_\perp/aT_c)^{1/4}$ and $\xi_{\parallel0}=(c_\parallel/aT_c)^{1/2}$ are the bare correlation lengths perpendicular and parallel to the plane, respectively. The static susceptibility $\chi(0)=(D_\perp q_*^4+|\Delta\tilde r|)^{-1}$ does not diverge at the unordered to modulated transition. Though it anticipates a transition with Curie--Weiss-like divergence at the lower point $a(T-T_c)=\Delta\tilde r=-D_\perp q_*^4<0$, this is cut off with a cusp at $\Delta\tilde r=0$. The elastic susceptibility, which is the reciprocal of the effective elastic modulus, is found in a similar way to the \op\ susceptibility: we must trace over $\eta$ and take the second variation of the resulting effective free energy functional of $\epsilon$ alone. Extremizing over $\eta$ yields \begin{equation} 0=\frac{\delta F[\eta,\epsilon]}{\delta\eta(x)}\bigg|_{\eta=\eta_\star} =\frac{\delta F_\op[\eta]}{\delta\eta(x)}\bigg|_{\eta=\eta_\star}-b\epsilon_\X(x), \label{eq:implicit.eta} \end{equation} which implicitly gives $\eta_\star[\epsilon]$, the \op\ conditioned on the configuration of the strain. Since $\eta_\star$ is a functional of $\epsilon_\X$ alone, only the modulus $C_\X$ will be modified from its bare value $C^0_\X$. Though the differential equation for $\eta_*$ cannot be solved explicitly, we can use the inverse function theorem to make use of it anyway. First, denote by $\eta_\star^{-1}[\eta]$ the inverse functional of $\eta_\star$ implied by \eqref{eq:implicit.eta}, which gives the function $\epsilon_\X$ corresponding to each solution of \eqref{eq:implicit.eta} it receives. This we can immediately identify from \eqref{eq:implicit.eta} as $\eta^{-1}_\star[\eta](x)=b^{-1}(\delta F_\op[\eta]/\delta\eta(x))$. Now, we use the inverse function theorem to relate the functional reciprocal of the derivative of $\eta_\star[\epsilon]$ with respect to $\epsilon_\X$ to the derivative of $\eta^{-1}_\star[\eta]$ with respect to $\eta$, yielding \begin{equation} \begin{aligned} \bigg(\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')}\bigg)^\recip &=\frac{\delta\eta_\star^{-1}[\eta](x)}{\delta\eta(x')}\bigg|_{\eta=\eta^*[\epsilon]} \\ &=b^{-1}\frac{\delta^2F_\op[\eta]}{\delta\eta(x)\delta\eta(x')}\bigg|_{\eta=\eta^*[\epsilon]}. \end{aligned} \label{eq:inv.func} \end{equation} Next, \eqref{eq:implicit.eta} and \eqref{eq:inv.func} can be used in concert with the ordinary rules of functional calculus to yield the second variation \begin{widetext} \begin{equation} \begin{aligned} \frac{\delta^2F[\eta_\star[\epsilon],\epsilon]}{\delta\epsilon_\X(x)\delta\epsilon_\X(x')} &=C^0_\X\delta(x-x')- 2b\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')} -b\int dx''\,\frac{\delta^2\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')\delta\epsilon_\X(x'')}\epsilon_\X(x'') \\ &+\int dx''\,\frac{\delta^2\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x)\delta\epsilon_\X(x')}\frac{\delta F_\op[\eta]}{\delta\eta(x'')}\bigg|_{\eta=\eta_\star[\epsilon]} +\int dx''\,dx'''\,\frac{\delta\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x)}\frac{\delta\eta_\star[\epsilon](x''')}{\delta\epsilon_\X(x')}\frac{\delta^2F_\op[\eta]}{\delta\eta(x'')\delta\eta(x''')}\bigg|_{\eta=\eta_\star[\epsilon]} \\ &=C^0_\X\delta(x-x')- 2b\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')} -b\int dx''\,\frac{\delta^2\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')\delta\epsilon_\X(x'')}\epsilon_\X(x'') \\ &+\int dx''\,\frac{\delta^2\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x)\delta\epsilon_\X(x')}(b\epsilon_\X(x'')) +b\int dx''\,dx'''\,\frac{\delta\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x)}\frac{\delta\eta_\star[\epsilon](x''')}{\delta\epsilon_\X(x')} \bigg(\frac{\partial\eta_\star[\epsilon](x'')}{\partial\epsilon_\X(x''')}\bigg)^\recip\\ &=C^0_\X\delta(x-x')- 2b\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')} +b\int dx''\,\delta(x-x'')\frac{\delta\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x')} =C^0_\X\delta(x-x')-b\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')}. \end{aligned} \label{eq:big.boy} \end{equation} \end{widetext} The elastic modulus is given by the second variation evaluated at the extremized strain $\langle\epsilon\rangle$. To calculate it, note that evaluating the second variation of $F_\op$ in \eqref{eq:inv.func} at $\langle\epsilon\rangle$ (or $\eta_\star(\langle\epsilon\rangle)=\langle\eta\rangle$) yields \begin{equation} \bigg(\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')}\bigg)^\recip\bigg|_{\epsilon=\langle\epsilon\rangle} =b^{-1}\chi^\recip(x,x')+\frac{b}{C^0_\X}\delta(x-x'), \label{eq:recip.deriv.op} \end{equation} where $\chi^\recip$ is the \op\ susceptibility given by \eqref{eq:sus_def}. Upon substitution into \eqref{eq:big.boy} and taking the Fourier transform of the result, we finally arrive at \begin{equation} C_\X(q) =C^0_\X-b\bigg(\frac1{b\chi(q)}+\frac b{C^0_\X}\bigg)^{-1} =C^0_\X\bigg(1+\frac{b^2}{C^0_\X}\chi(q)\bigg)^{-1}. \label{eq:elastic.susceptibility} \end{equation} Though not relevant here, this result generalizes to multicomponent \op s. What does \eqref{eq:elastic.susceptibility} predict in the vicinity of the \ho\ transition? Near the disordered to modulated transition, the static modulus is given by \begin{equation} C_\X(0)=C_\X^0\bigg[1+\frac{b^2}{C_\X^0}\big(D_\perp q_*^4+|\Delta\tilde r|\big)^{-1}\bigg]^{-1}. \label{eq:static_modulus} \end{equation} This corresponds to a softening in the $\X$-modulus at the transition that is cut off with a cusp of the form $|\Delta\tilde r|^\gamma\propto|T-T_c|^\gamma$ with $\gamma=1$. This is our main result. The only \op\ irreps that couple linearly with strain and reproduce the topology of the \urusi\ phase diagram are $\Bog$ and $\Btg$. For either of these irreps, the transition into a modulated rather than uniform phase masks traditional signatures of a continuous transition by locating thermodynamic singularities at nonzero $q=q_*$. The remaining clue at $q=0$ is a particular kink in the corresponding modulus. \begin{figure}[htpb] \centering \includegraphics[width=\columnwidth]{fig-stiffnesses} \caption{ \Rus\ measurements of the elastic moduli of \urusi\ as a function of temperature from \cite{ghosh_single-component_nodate} (blue, solid) alongside fits to theory (magenta, dashed). The solid yellow region shows the location of the \ho\ phase. (a) $\Btg$ modulus data and fit to standard form \cite{varshni_temperature_1970}. (b) $\Bog$ modulus data and fit to \eqref{eq:static_modulus}. The fit gives $C^0_\Bog\simeq\big[71-(0.010\,\K^{-1})T\big]\,\GPa$, $D_\perp q_*^4/b^2\simeq0.16\,\GPa^{-1}$, and $a/b^2\simeq6.1\times10^{-4}\,\GPa^{-1}\,\K^{-1}$. Addition of an additional parameter to fit the standard bare modulus \cite{varshni_temperature_1970} led to poorly constrained fits. (c) $\Bog$ modulus data and fit of \emph{bare} $\Bog$ modulus. (d) $\Bog$ modulus data and fit transformed by $[C^0_\Bog(C^0_\Bog/C_\Bog-1)]]^{-1}$, which is predicted from \eqref{eq:static_modulus} to equal $D_\perp q_*^4/b^2+a/b^2|T-T_c|$, e.g., an absolute value function. The failure of the Ginzburg--Landau prediction below the transition is expected on the grounds that the \op\ is too large for the free energy expansion to be valid by the time the Ginzburg temperature is reached. } \label{fig:data} \end{figure} \emph{Comparison to experiment.} \Rus\ experiments \cite{ghosh_single-component_nodate} yield the individual elastic moduli broken into irrep symmetries; the $\Bog$ and $\Btg$ components defined in \eqref{eq:strain-components} are shown in Figures \ref{fig:data}(a--b). The $\Btg$ modulus doesn't appear to have any response to the presence of the transition, exhibiting the expected linear stiffening upon cooling from room temperature, with a low-temperature cutoff at some fraction of the Debye temperature \cite{varshni_temperature_1970}. The $\Bog$ modulus has a dramatic response, softening over the course of roughly $100\,\K$ and then cusping at the \ho\ transition. While the low-temperature response is not as dramatic as the theory predicts, mean field theory---which is based on a small-$\eta$ expansion---will not work quantitatively far below the transition where $\eta$ has a large nonzero value and higher powers in the free energy become important. The data in the high-temperature phase can be fit to the theory \eqref{eq:static_modulus}, with a linear background modulus $C^0_\Bog$ and $\tilde r-\tilde r_c=a(T-T_c)$, and the result is shown in Figure \ref{fig:data}(b). The data and theory appear quantitatively consistent in the high temperature phase, suggesting that \ho\ can be described as a $\Bog$-nematic phase that is modulated at finite $q$ along the $c-$axis. The predicted softening appears over hundreds of Kelvin; Figures \ref{fig:data}(c--d) show the background modulus $C_\Bog^0$ and the \op--induced response isolated from each other. We have seen that the mean-field theory of a $\Bog$ \op\ recreates the topology of the \ho\ phase diagram and the temperature dependence of the $\Bog$ elastic modulus at zero pressure. This theory has several other physical implications. First, the association of a modulated $\Bog$ order with the \ho\ phase implies a \emph{uniform} $\Bog$ order associated with the \afm\ phase, and moreover a uniform $\Bog$ strain of magnitude $\langle\epsilon_\Bog\rangle^2=b^2\tilde r/4u(C^0_\Bog)^2$, which corresponds to an orthorhombic structural phase. Orthorhombic symmetry breaking was recently detected in the \afm\ phase of \urusi\ using x-ray diffraction, a further consistency of this theory with the phenomenology of \urusi\ \cite{choi_pressure-induced_2018}. Second, as the Lifshitz point is approached from low pressure, this theory predicts that the modulation wavevector $q_*$ should vanish continuously. Far from the Lifshitz point we expect the wavevector to lock into values commensurate with the space group of the lattice, and moreover that at zero pressure, where the \rus\ data here was collected, the half-wavelength of the modulation should be commensurate with the lattice spacing $a_3\simeq9.68\,\A$, or $q_*=\pi/a_3\simeq0.328\,\A^{-1}$ \cite{meng_imaging_2013, broholm_magnetic_1991, wiebe_gapped_2007, bourdarot_precise_2010}. In between these two regimes, the ordering wavevector should shrink by jumping between ever-closer commensurate values in the style of the devil's staircase \cite{bak_commensurate_1982}. This motivates future \rus\ experiments done at pressure, where the depth of the cusp in the $\Bog$ modulus should deepen (perhaps with these commensurability jumps) at low pressure and approach zero like $q_*^4\sim(c_\perp/2D_\perp)^2$ near the Lifshitz point. {\color{blue} Moreover, } \brad{Should also motivate x-ray and neutron-diffraction experiments to look for new q's - mentioning this is important if we want to get others interested, no one else does RUS...} Alternatively, \rus\ done at ambient pressure might examine the heavy fermi liquid to \afm\ transition by doping. \brad{We have to be careful, someone did do some doping studies and it's not clear exctly what's going on}. The presence of spatial commensurability known to be irrelevant to critical behavior at a one-component disordered to modulated transition, and therefore is not expected to modify the thermodynamic behavior otherwise \cite{garel_commensurability_1976}. There are two apparent discrepancies between the orthorhombic strain in the phase diagram presented by \cite{choi_pressure-induced_2018} and that predicted by our mean field theory. The first is the apparent onset of the orthorhombic phase in the \ho\ state prior to the onset of \afm. As \cite{choi_pressure-induced_2018} notes, this could be due to the lack of an ambient pressure calibration for the lattice constant. The second discrepancy is the onset of orthorhombicity at higher temperatures than the onset of \afm. Susceptibility data sees no trace of another phase transition at these higher temperatures \cite{inoue_high-field_2001}. We suspect that the high-temperature orthorhombic signature is not the result of a bulk phase, and could be due to the high energy (small-wavelength) nature of x-rays as an experimental probe: \op\ fluctuations should lead to the formation of orthorhombic regions on the order of the correlation length that become larger and more persistent as the transition is approached. Three dimensions is below the upper critical dimension $4\frac12$ of a one-component disordered to modulated transition, and so mean field theory should break down sufficiently close to the critical point due to fluctuations, at the Ginzburg temperature \cite{hornreich_lifshitz_1980, ginzburg_remarks_1961}. Magnetic phase transitions tend to have Ginzburg temperature of order one. Our fit above gives $\xi_{\perp0}q_*=(D_\perp q_*^4/aT_c)^{1/4}\simeq2$, which combined with the speculation of $q_*\simeq\pi/a_3$ puts the bare correlation length $\xi_{\perp0}$ at about what one would expect for a generic magnetic transition. The agreement of this data in the $t\sim0.1$--10 range with the mean field exponent suggests that this region is outside the Ginzburg region, but an experiment may begin to see deviations from mean field behavior within around several degrees Kelvin of the critical point. A \rus\ experiment with more precise temperature resolution near the critical point may be able to resolve a modified cusp exponent $\gamma\simeq1.31$ \cite{guida_critical_1998}, since the universality class of a uniaxial modulated one-component \op\ is $\mathrm O(2)$ \cite{garel_commensurability_1976}. We should not expect any quantitative agreement between mean field theory and experiment in the low temperature phase since, by the point the Ginzburg criterion is satisfied, $\eta$ is order one and the Landau--Ginzburg free energy expansion is no longer valid. \emph{Conclusion and Outlook.} We have developed a general phenomenological treatment of \ho\ \op s with the potential for linear coupling to strain. The two representations with mean field phase diagrams that are consistent with the phase diagram of \urusi\ are $\Bog$ and $\Btg$. Of these, only a staggered $\Bog$ \op is consistent with zero-pressure \rus\ data, with a cusp appearing in the associated elastic modulus. In this picture, the \ho\ phase is characterized by uniaxial modulated $\Bog$ order, while the \afm\ phase is characterized by uniform $\Bog$ order. \brad{We need to be a bit more explicit about what we think is going on with \afm - is it just a parasitic phase? Is our modulated phase somehow "moduluated \afm" (can you modualte AFM in such as way as to make it disappear? Some combination of orbitals?)} The corresponding prediction of uniform $\Bog$ symmetry breaking in the \afm\ phase is consistent with recent diffraction experiments \cite{choi_pressure-induced_2018}, except for the apparent earlier onset in temperature of the $\Bog$ symmetry breaking than AFM, which we believe to be due to fluctuating order above the actual phase transition. This work motivates both further theoretical work regarding a microscopic theory with modulated $\Bog$ order, and preforming \rus\ experiments at pressure that could further support or falsify this idea. \begin{acknowledgements} This research was supported by NSF DMR-1719490 and DMR-1719875. \end{acknowledgements} \bibliographystyle{apsrev4-1} \bibliography{hidden_order} \end{document}