2 Interpretation of stability analysis around a turbulent mean flow

Mathematical interpretation of the stability analysis around the mean flow is not as straightforward as the stability analysis around a steady solution of the Navier–Stokes equations. Nevertheless, a qualitative mathematical and physical interpretation of mean flow stability results and qualitative criteria for their validity can be found. The argument below follows the main lines presented in Turton, Tuckerman & Barkley (Reference Turton, Tuckerman and Barkley2015). In the present study, a triple decomposition of the flow field is introduced following Reynolds & Hussain (Reference Reynolds and Hussain1972):

(2.1) $$\begin{eqnarray}\boldsymbol{u}=\overline{\boldsymbol{u}}+\widetilde{\boldsymbol{u}}+\boldsymbol{u}^{\prime },\end{eqnarray}$$

where $\overline{\phantom{u}}$ is the time-average operator, $(\,\widetilde{\phantom{u}}+\overline{\phantom{u}})$ is the phase-average operator, and $\boldsymbol{u}^{\prime }=\boldsymbol{u}-\overline{\boldsymbol{u}}-\widetilde{\boldsymbol{u}}$ is the fluctuation with zero phase average. The three terms are the mean flow ( $\overline{\boldsymbol{u}}$ ), the organized wave containing all coherent time-periodic large-scale motions ( $\widetilde{\boldsymbol{u}}$ ), and the stochastic part containing the remaining incoherent turbulent motions ( $\boldsymbol{u}^{\prime }$ ). The equation for the mean flow is obtained by taking the time average of the Navier–Stokes equations:

(2.2) $$\begin{eqnarray}\overline{\boldsymbol{U}}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\overline{\boldsymbol{U}}=-\boldsymbol{{\rm\nabla}}\overline{P}+\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }(Re^{-1}\overline{\unicode[STIX]{x1D64E}}-\overline{\widetilde{\boldsymbol{u}}\widetilde{\boldsymbol{u}}}-\overline{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}),\end{eqnarray}$$

while the organized wave satisfies the phase-averaged Navier–Stokes equations, with (2.2) subtracted:

(2.3) $$\begin{eqnarray}\frac{\partial \widetilde{\boldsymbol{u}}}{\partial t}+\overline{\boldsymbol{U}}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\widetilde{\boldsymbol{u}}+\widetilde{\boldsymbol{u}}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\overline{\boldsymbol{U}}=-\boldsymbol{{\rm\nabla}}\widetilde{p}+\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }(Re^{-1}\widetilde{\unicode[STIX]{x1D668}}-\widetilde{\widetilde{\boldsymbol{u}}\widetilde{\boldsymbol{u}}}-\widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}).\end{eqnarray}$$

In the above, $\unicode[STIX]{x1D64E}=\boldsymbol{{\rm\nabla}}\boldsymbol{U}+\boldsymbol{{\rm\nabla}}\boldsymbol{U}^{\text{T}}$ is the mean flow shear stress tensor, and $\widetilde{\unicode[STIX]{x1D668}}$ the stress tensor of the organized wave. We will proceed by assuming that the coherent motions consist of discrete fundamental limit cycles and their harmonics, and can hence be Fourier-decomposed as: $\widetilde{\boldsymbol{u}}\approx \sum _{m>0}\sum _{n\neq 0}\boldsymbol{u}_{m,n}\exp (\text{i}n{\it\omega}_{m}t)$ .

For simplicity, let us first consider the case where limit cycles with different $m$ , and their products, are harmonically unrelated to each other (at the end of the section, we will return to the case where they are harmonically related). When substituting the Fourier decomposition of the coherent part into (2.2), we obtain:

(2.4) $$\begin{eqnarray}\overline{\boldsymbol{U}}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\overline{\boldsymbol{U}}=-\boldsymbol{{\rm\nabla}}\overline{P}+\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }\left(Re^{-1}\overline{\unicode[STIX]{x1D64E}}-\mathop{\sum }_{m>0}\mathop{\sum }_{n>0}\boldsymbol{u}_{m,n}\boldsymbol{u}_{m,-n}-\overline{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}\right).\end{eqnarray}$$

This shows that the mean flow is influenced by the coherent motions, through the interaction of each fundamental mode with its conjugate, and the interaction of each harmonic with its conjugate. The mean flow is also influenced by the Reynolds stresses arising from the incoherent motions.

Similarly, when substituting the Fourier decomposition into (2.3), we obtain for $n=1$ (the limit cycle fundamental):

(2.5) $$\begin{eqnarray}-\text{i}{\it\omega}\boldsymbol{u}_{m,1}+\overline{\boldsymbol{U}}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\boldsymbol{u}_{m,1}+\boldsymbol{u}_{m,1}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\overline{\boldsymbol{U}}=-\boldsymbol{{\rm\nabla}}p_{m,1}+\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }\left(\!Re^{-1}\unicode[STIX]{x1D64E}_{m,1}-\!\mathop{\sum }_{n\neq 0,1}\!\boldsymbol{u}_{m,n}\boldsymbol{u}_{m,1-n}-(\widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }})_{m}\!\right)\!.\end{eqnarray}$$

Hence, the limit cycle $m$ may be influenced by the coherent motions, through the interaction of the first harmonic $\boldsymbol{u}_{m,2}$ with the conjugate of the fundamental $\boldsymbol{u}_{m,-1}$ , and the interaction of each higher harmonic with the conjugate of its preceding harmonic. It may also be influenced by $(\widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }})_{m}$ , which is the oscillation of the incoherent Reynolds stresses at frequency ${\it\omega}={\it\omega}_{m}$ .

Let us now introduce the linearized Navier–Stokes operator, where the linearization is performed around the mean flow, acting on any velocity field $\boldsymbol{u}$ :

(2.6) $$\begin{eqnarray}{\mathcal{L}}_{\overline{U}}(\boldsymbol{u})=\overline{\boldsymbol{U}}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\boldsymbol{u}+\boldsymbol{u}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\overline{\boldsymbol{U}}+\boldsymbol{{\rm\nabla}}p-\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }(Re^{-1}\boldsymbol{u}).\end{eqnarray}$$

Equation (2.5) can be formally written as:

(2.7) $$\begin{eqnarray}{\mathcal{L}}_{\overline{U}}(\boldsymbol{u}_{m,1})=\text{i}{\it\omega}_{m}\boldsymbol{u}_{m,1}-{\mathcal{N}}_{1}-\widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}.\end{eqnarray}$$

Here, ${\mathcal{N}}_{1}$ is a nonlinear harmonic interaction term given by:

(2.8) $$\begin{eqnarray}{\mathcal{N}}_{1}=\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }(\boldsymbol{u}_{m,2}\boldsymbol{u}_{m,-1}+\boldsymbol{u}_{m,-1}\boldsymbol{u}_{m,2}+\boldsymbol{u}_{m,3}\boldsymbol{u}_{m,-2}+\boldsymbol{u}_{m,-2}\boldsymbol{u}_{m,3}+\cdots \,).\end{eqnarray}$$

No assumptions have been introduced so far, apart from the coherent motions being discrete (and this assumption could be relaxed by writing the coherent motions as an integral instead of a sum). It can be seen that (2.5) forms a linear eigenvalue problem for the fundamental mode $\boldsymbol{u}_{m,1}$ if and only if either of the two options is true:

1. (i) ${\mathcal{N}}_{1}+\widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}=0$ .

As noted by Turton et al. (Reference Turton, Tuckerman and Barkley2015), we observe that the fundamental mode $m$ is then an exact eigenmode of the Navier–Stokes operator linearized around the mean flow:

(2.9) $$\begin{eqnarray}{\mathcal{L}}_{\overline{U}}(\boldsymbol{u}_{m,1})=\text{i}{\it\omega}\boldsymbol{u}_{m,1},\end{eqnarray}$$

with the eigenvalue ${\it\omega}$ which has zero growth rate and the frequency of the fundamental limit cycle.

1. (ii) ${\mathcal{N}}_{1}+\widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}={\mathcal{A}}\boldsymbol{u}_{m,1}$

where ${\mathcal{A}}$ is a linear operator. Then, the fundamental mode $m$ is then an exact eigenmode of the modified Navier–Stokes operator $({\mathcal{L}}_{\overline{U}}+{\mathcal{A}})$ :

(2.10) $$\begin{eqnarray}({\mathcal{L}}_{\overline{U}}+{\mathcal{A}})\{\boldsymbol{u}_{m,1}\}=\text{i}{\it\omega}\boldsymbol{u}_{m,1}.\end{eqnarray}$$

Again, the fundamental mode $m$ will then have zero growth rate, and the same frequency as the limit cycle. (The first option is actually a special case of the second one, obtained where ${\mathcal{A}}=0$ .)

To relate the above constraints into qualitative properties of a flow model, we proceed similarly to Turton et al. (Reference Turton, Tuckerman and Barkley2015), who pointed out that if the amplitude of the fundamental mode is $\Vert \boldsymbol{u}_{m,1}\Vert ={\it\epsilon}$ , the harmonics often decay as $\boldsymbol{u}_{m,n}\propto O({\it\epsilon}^{n})$ . If this is the case, then ${\mathcal{N}}_{1}=O({\it\epsilon}^{3})$ (where the first term of ${\mathcal{N}}_{1}$ is the largest: $\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }(\boldsymbol{u}_{m,2}\boldsymbol{u}_{m,-1})=O({\it\epsilon}^{3})$ ), and if $A\boldsymbol{u}_{m,1}=\widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}$ , then:

(2.11) $$\begin{eqnarray}({\mathcal{L}}+{\mathcal{A}})\{\boldsymbol{u}_{m,1}\}-\text{i}{\it\omega}\boldsymbol{u}_{m,1}=O({\it\epsilon}^{3}).\end{eqnarray}$$

Since the right hand side terms are $O({\it\epsilon}^{3})$ , the fundamental limit cycle is still approximately an eigenmode of $({\mathcal{L}}+{\mathcal{A}})$ . In particular this assumption approximately implies that the second harmonic needs to be an order of magnitude weaker than the fundamental. A similar argument based on the relative amplitude of the second harmonic has been emphasized by several authors, for example Sipp & Lebedev (Reference Sipp and Lebedev2007). In the present study, we have verified a posteriori that the second harmonic is invisible among the broadband turbulent spectrum, and hence we conclude similarly to Turton et al. (Reference Turton, Tuckerman and Barkley2015) that ${\mathcal{N}}_{1}\leqslant O({\it\epsilon}^{2})$ in this flow.

Now, for mean flow stability to reproduce the fundamental limit cycle, the broadband turbulent motions still need to satisfy ${\mathcal{A}}\boldsymbol{u}_{m,1}=\widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}$ . In this study, we will assume the eddy-viscosity hypothesis for the incoherent motions, such that: ${\it\mu}_{t}(\unicode[STIX]{x1D64E}_{m,1}^{\star })=\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }(\widetilde{\boldsymbol{u}^{\star ^{\prime }}\boldsymbol{u}^{\star ^{\prime }}})$ , where ${\it\mu}_{t}$ denotes turbulent viscosity and the stars denote dimensional variables.

Finally, in the case that the frequencies of other limit cycles are harmonically related to the limit cycle under study, their amplitudes also need to be an order of magnitude smaller than that of the dominant limit cycle under investigation (and their harmonics need to decay as rapidly as for the dominant mode). If a product of two fundamental limit cycles ( $m=i$ and $m=j$ ) has a frequency equal to the dominant mode ( $m=1$ ), i.e. if ${\it\omega}_{1}={\it\omega}_{i}\pm {\it\omega}_{j}$ , then they may contribute to (2.5) but are by definition ${\leqslant}O({\it\epsilon}^{2})$ . The sum of such modes needs to converge rapidly enough to remain $O({\it\epsilon}^{2})$ .

2.3 Relation to mean-field theory and weakly nonlinear stability approaches

The above approach can be related to the mean-field theory by Stuart (Reference Stuart1958, Reference Stuart1971), used and developed in a long line of studies for model reduction. A Galerkin projection of the Navier–Stokes equations formally written as:

(2.12) $$\begin{eqnarray}\frac{\text{d}}{\text{d}t}a_{i}=\frac{1}{Re}\mathop{\sum }_{j=0}^{N+1}l_{ij}a_{j}+\mathop{\sum }_{j=0}^{N+1}q_{ijk}a_{j}a_{k},\end{eqnarray}$$

where $i>0$ , $l_{ij}$ is the linear operator part of Navier–Stokes, $q_{ijk}$ the nonlinear operator part originating from the advection term and $a_{i}$ are constant coefficients. Here, the mean flow is the zeroth mode:

(2.13) $$\begin{eqnarray}\boldsymbol{u}_{0}=\overline{\boldsymbol{U}}.\end{eqnarray}$$

The mode number $N+1$ is the zero-frequency shift mode, representing the effect of coherent Reynolds stresses on the mean flow. (In the special case of a flow saturating towards a limit cycle starting from a steady base flow, the shift mode can be described as the difference between the period-averaged mean flow and the base flow: $\overline{\boldsymbol{U}}=\boldsymbol{U}_{s}+\boldsymbol{u}_{{\it\Delta}}$ . Where $\overline{\boldsymbol{U}}$ is the mean flow, $\boldsymbol{U}_{s}$ the steady solution to Navier–Stokes equations.) The minimal Galerkin system for a simple limit cycle is obtained with  $N=2$ :

(2.14) $$\begin{eqnarray}\boldsymbol{u}=\boldsymbol{u}_{0}+\mathop{\sum }_{i=1}^{2}a_{i}\boldsymbol{u}_{i}+a_{{\it\Delta}}\boldsymbol{u}_{{\it\Delta}},\end{eqnarray}$$

where the modes 1–2 are the real and imaginary parts of the limit cycle (obtained in Noack et al. (Reference Noack, Afanasiev, Morzynski, Tadmor and Thiele2003) from the leading mode pair in a POD decomposition of the saturated state), and $\boldsymbol{u}_{{\it\Delta}}$ is the shift mode. The system is further simplified(Noack et al. Reference Noack, Afanasiev, Morzynski, Tadmor and Thiele2003) by a Kryloff–Bogoliubov ansatz to yield the amplitudes:

(2.15) $$\begin{eqnarray}\displaystyle & \displaystyle a_{1}=A({\it\epsilon}t)\cos ({\it\omega}t), & \displaystyle\end{eqnarray}$$

(2.16) $$\begin{eqnarray}\displaystyle & \displaystyle a_{2}=A({\it\epsilon}t)\sin ({\it\omega}t), & \displaystyle\end{eqnarray}$$

(2.17) $$\begin{eqnarray}\displaystyle & \displaystyle a_{{\it\Delta}}=B({\it\epsilon}t), & \displaystyle\end{eqnarray}$$

where ${\it\epsilon}$ is a slow time scale much longer than the limit cycle oscillation period. This minimal Galerkin system for a simple limit cycle reproduces the qualitative behaviour of the cylinder wake, such as saturation to the limit cycle from steady state, and influence of stabilizing control (Noack et al. Reference Noack, Afanasiev, Morzynski, Tadmor and Thiele2003). Relating this to (2.4), we can interpret the shift mode as a change of the mean flow due to a change in the amplitude of the coherent fluctuation, through the term: $\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }(\overline{\boldsymbol{u}_{1,1}\boldsymbol{u}_{1,-1}})$ . The assumptions behind the minimal system are essentially the same as in the present study. The energy transfer from higher harmonics to the fundamental mode is ignored, while the energy transfer from the fundamental mode to the mean flow is taken into account. Similarly to the shift mode, a one-way effect of high-frequency actuation on Reynolds stresses has also been incorporated (Luchtenburg et al. Reference Luchtenburg, Günther, Noack, King and Tadmor2009), and as in the present study the effect of turbulence using different eddy-viscosity models (Östh et al. Reference Östh, Noack, Krajnovic, Barros and Bore2014; Protas et al. Reference Protas, Noack and Östh2015). In some cases, a priori criteria for validity of the model may be found. Sipp & Lebedev (Reference Sipp and Lebedev2007) formulated a weakly nonlinear analysis of flows near the critical Reynolds number for bifurcation, and formulated criteria for validity of mean flow stability analysis based on Landau coefficients ${\it\mu}$ and ${\it\nu}$ . Of these, ${\it\mu}$ represents the magnitude of the interaction between the fundamental eigenmode and the zeroth harmonic (i.e. the mean flow change induced by the eigenmode, which can be compared to the shift mode), while ${\it\nu}$ was the magnitude of the interaction between the eigenmode and its first harmonic. Consistently with the other models, ${\it\nu}\ll {\it\mu}$ (small relative amplitude of the harmonic) indicated good behaviour of the mean flow stability. However, the criteria for frequency and growth rate were not the same. The mean flow stability would return a marginally stable mode if the ratio of imaginary parts ( ${\it\mu}_{i}/{\it\lambda}_{i}$ ) was small, while the frequency of the limit cycle would be well approximated if the ratio of the real parts ( ${\it\mu}_{r}/{\it\lambda}_{r}$ ) was small. This explains the observation that in many mean flow analyses the frequency and shape of the limit cycle is well reproduced, while the growth rate may remain strictly positive (especially when eddy-viscosity models are not used).

3 Problem definition

The geometry consists of an inner pipe (without flow), and an outer coaxial inlet called ‘the swirler’ in this manuscript. Two non-dimensional parameters define the characteristics of the flow: the Reynolds number $Re$ and the swirl ratio $S_{w}$ . The Reynolds number is defined as

(3.1) $$\begin{eqnarray}Re=\frac{U_{e}D}{{\it\nu}},\end{eqnarray}$$

where $U_{e}$ is the mean velocity at the swirler exit ( $U_{e}=Q/A$ , where $Q$ is the flow rate and $A$ the swirler exit area), and $D$ is the outer diameter of the swirler at the exit (these scales are illustrated in figure 1). The swirl ratio is defined as:

(3.2) $$\begin{eqnarray}S_{w}=\frac{W_{e}}{U_{e}},\end{eqnarray}$$

where $W_{e}$ is the mean azimuthal velocity at the swirler exit. By these definitions, the non-dimensional parameters become $Re=4800$ and $Sw=1.1$ . The coordinate system is defined in figure 1.

3.1 Eddy-viscosity model

To proceed from (2.5) to create an eddy-viscosity model for the incoherent fluctuations, we need to set $\widetilde{\widetilde{\boldsymbol{u}}\widetilde{\boldsymbol{u}}}\approx 0$ and $\overline{\widetilde{\boldsymbol{u}}\widetilde{\boldsymbol{u}}}=0$ . We can now define an eddy-viscosity field for both the mean flow and the organized wave by the use of the following Boussinesq relations between Reynolds stresses and the turbulent viscosity:

(3.3) $$\begin{eqnarray}\displaystyle & \displaystyle \overline{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}={\textstyle \frac{2}{3}}\overline{k}\unicode[STIX]{x1D644}-2\overline{{\it\nu}}_{t}\overline{\unicode[STIX]{x1D64E}} & \displaystyle\end{eqnarray}$$

(3.4) $$\begin{eqnarray}\displaystyle & \displaystyle \widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}={\textstyle \frac{2}{3}}\tilde{k}\unicode[STIX]{x1D644}-2\overline{{\it\nu}}_{t}\tilde{\unicode[STIX]{x1D64E}}-2\tilde{{\it\nu}}_{t}\overline{\unicode[STIX]{x1D64E}}, & \displaystyle\end{eqnarray}$$

where $k=\boldsymbol{u}\boldsymbol{\cdot }\boldsymbol{u}$ is the kinetic energy of the organized wave, and $\unicode[STIX]{x1D644}$ is the identity tensor. We now assume that the eddy-viscosity field itself is not oscillated by the perturbation, $\widetilde{{\it\nu}}_{t}=0$ , and similarly for the turbulent kinetic energy: $\widetilde{k}=0$ , to obtain:

(3.5) $$\begin{eqnarray}\displaystyle & \overline{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}={\textstyle \frac{2}{3}}\overline{k}\unicode[STIX]{x1D644}-2\overline{{\it\nu}}_{t}\overline{\unicode[STIX]{x1D64E}} & \displaystyle\end{eqnarray}$$

(3.6) $$\begin{eqnarray}\displaystyle & \widetilde{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}=-2\overline{{\it\nu}}_{t}\widetilde{\unicode[STIX]{x1D64E}}. & \displaystyle\end{eqnarray}$$

This means that the eddy viscosity $\overline{{\it\nu}_{t}}$ can be determined from the averages using (3.5), and used as is for the oscillating Reynolds stresses (3.6). This is called the Newtonian eddy model. Looking carefully, the Newtonian eddy model is slightly inconsistent in that the mean flow averages will always contain the coherent fluctuations. Hence, it is strictly valid only when $\widetilde{\widetilde{\boldsymbol{u}}\widetilde{\boldsymbol{u}}}\approx 0$ and $\overline{\widetilde{\boldsymbol{u}}\widetilde{\boldsymbol{u}}}=0$ are both very small. This model was pointed out by Reynolds & Hussain (Reference Reynolds and Hussain1972) to work best for ‘relatively low frequency weak oscillations with wavelength considerably larger than the dominant scales of turbulence’.

To determine $\overline{{\it\nu}_{t}}$ , we take the Frobenius product between (3.5) and $\unicode[STIX]{x1D64E}$ , yielding:

(3.7) $$\begin{eqnarray}\overline{{\it\nu}}_{t}=-\frac{\overline{\boldsymbol{u}^{\prime }\boldsymbol{u}^{\prime }}\boldsymbol{ : }\overline{\unicode[STIX]{x1D64E}}}{2\overline{\unicode[STIX]{x1D64E}}\boldsymbol{ : }\overline{\unicode[STIX]{x1D64E}}},\end{eqnarray}$$

where $\boldsymbol{ : }$ is the Frobenius product, defined in Cartesian tensor notation by $\unicode[STIX]{x1D63C}\boldsymbol{ : }\unicode[STIX]{x1D63D}=\unicode[STIX]{x1D608}_{ij}\unicode[STIX]{x1D609}_{ij}$ .

Finally, we obtain the modified linear stability equation:

(3.8) $$\begin{eqnarray}{\it\sigma}\boldsymbol{u}+\overline{\boldsymbol{U}}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\boldsymbol{u}+\boldsymbol{u}\boldsymbol{\cdot }\boldsymbol{{\rm\nabla}}\overline{\boldsymbol{U}}=-\boldsymbol{{\rm\nabla}}p+\boldsymbol{{\rm\nabla}}\boldsymbol{\cdot }[Re_{eff}^{-1}(\boldsymbol{{\rm\nabla}}\boldsymbol{u}+\boldsymbol{{\rm\nabla}}\boldsymbol{u}^{\text{T}})],\end{eqnarray}$$

where

(3.9) $$\begin{eqnarray}Re_{eff}=\frac{{\it\nu}}{({\it\nu}+\overline{{\it\nu}}_{t})}Re,\end{eqnarray}$$

and the complex frequency will be denoted by ${\it\sigma}=-\text{i}{\it\omega}$ in the rest of this paper.

We note that a different approach could have been applied here, by performing the whole analysis around a turbulence model equation (‘base flow approach’), such as Unsteady RANS with the Spalart–Allmaras model as in Meliga et al. (Reference Meliga, Pujals and Serre2012b ). In that case, the base flow would be a fixed point of the URANS equations, which approximates the mean flow within the limit of validity of the turbulence model. The turbulence model (e.g. Spalart–Allmaras) would need to be linearized around the fixed point. This approach is mathematically fully consistent, and could be interesting to attempt in a future study. There is, however, a reason to believe that in swirling flow like the present one, the results from URANS might not be accurate due to a bad estimation of the mean flow swirl profile. A mathematically and physically fully consistent approach would be to linearize an algebraic Reynolds stress model (Wallin & Johansson Reference Wallin and Johansson2000), which is very complicated even in 1D (Gupta Reference Gupta2014), and was therefore considered to be out of the scope of the present study.

3.2 Linear global modes

Exploiting the homogeneity of the mean flow in the azimuthal direction, the perturbation $q$ takes the form:

(3.10a,b ) $$\begin{eqnarray}\boldsymbol{u}(z,r,{\it\theta},t)=\hat{\boldsymbol{u}}(z,r)\exp ({\it\sigma}t+\text{i}m{\it\theta}),\quad p(z,r,{\it\theta},t)=\hat{p}(z,r)\exp ({\it\sigma}t+\text{i}m{\it\theta}).\end{eqnarray}$$

Given an azimuthal wavenumber $m$ , (3.8) constitutes a linearized eigenvalue problem with the complex eigenvalue ${\it\sigma}$ . The imaginary part of the eigenvalue, ${\it\sigma}_{i}$ , is the angular oscillation frequency of the global mode, and the real part, ${\it\sigma}_{r}$ , gives the growth rate (usually called amplification rate in mean flow analysis). The Strouhal number is obtained from $St={\it\sigma}_{i}/2{\rm\pi}$ . The growth rate does not have a straightforward physical meaning when computing the modes around a mean flow. Because an oscillatory instability leads to a constant amplitude limit cycle around the mean, a close to neutral (zero) growth rate is expected for an oscillator in a mean flow analysis (Noack et al. Reference Noack, Afanasiev, Morzynski, Tadmor and Thiele2003). For the stability analysis, we set a zero Dirichlet velocity boundary condition for all boundaries.

The governing equations for the corresponding adjoint eigenmodes ( $\boldsymbol{u}^{+}$ , $p^{+}$ ) can be derived for example by using the Lagrange identity (Giannetti & Luchini Reference Giannetti and Luchini2007). In this study, the continuous adjoint equations of (3.8) with varying $Re_{eff}$ and (3.10) have not been explicitly derived. Instead, a so-called discrete adjoint approach is utilized (Schmid & Henningson Reference Schmid and Henningson2001), where the adjoint equations are numerically derived from the discretized matrix form of (3.8). Hence, no separate boundary conditions are set for the adjoint. For the molecular viscosity cases, a continuous adjoint formulation with zero Dirichlet boundary conditions is used, and verified against existing adjoint formulation in Nek5000 (appendix A). In both cases, the adjoint is normalized to satisfy: $\int _{V}\boldsymbol{u}^{+\ast }\boldsymbol{\cdot }\boldsymbol{u}\,\text{d}V=1$ , where $^{\ast }$ denotes the complex conjugate and $V$ the volume of the computational domain. Finally, the structural sensitivity is defined as the region where a local perturbation of the equation system results in the largest drift of the eigenvalue, and is given by $|\boldsymbol{u}||\boldsymbol{u}^{+}|$ . Here, the structural sensitivity is interpreted as the core of the instability or wavemaker (Giannetti & Luchini Reference Giannetti and Luchini2007).

4 Numerical methods

Two numerical codes have been used in this study. The Nek5000 code (Fischer Reference Fischer1997; Fischer, Lottes & Kerkemeier Reference Fischer, Lottes and Kerkemeier2008) was used for time integration of the nonlinear Navier–Stokes equations, which generated the DNS results and the POD modes presented in § 5.1. The global mode results without eddy viscosity included in appendix A were also obtained using Nek5000. The global mode results in the bulk of the manuscript were obtained by using the finite element package FreeFem++ (Hecht Reference Hecht2012). The variational formulation of the direct global mode equations including variable viscosity was derived and implemented in this work.

4.1 Nonlinear simulations and POD

Nek5000 is based on a spectral element method (Maday & Patera Reference Maday, Patera and Noor1989), which combines the accuracy of spectral methods with the flexibility of finite element methods. For details about the code implementation see Maday & Patera (Reference Maday, Patera and Noor1989). The same implementation including the Arnoldi method for modes in appendix A was used also in Lashgari et al. (Reference Lashgari, Tammisola, Citro, Brandt and Juniper2014).

In the DNS, the nonlinear Navier–Stokes equations were integrated forward in time for 481 non-dimensional time units, corresponding to around 40 flow through times from the nozzle inlet to the chamber exit. This simulation was run on high-performance computing clusters using between 256 and 1024 cores in parallel, and required over 100 000 CPU hours to complete. The time integration was performed by an explicit second-order extrapolation for the nonlinear terms, and an implicit second-order backwards-differentiation for the viscous terms, as in previous turbulent diffuser studies in Nek5000 (Ohlsson et al. Reference Ohlsson, Schlatter, Fischer and Henningson2010). The non-dimensional time step was kept at ${\rm\Delta}t=1\times 10^{-4}$ to satisfy the Courant–Friedrichs–Lewy (CFL) condition. Our accuracy in time integration should compare well with other turbulent flow studies with spectral element methods, for example Ohlsson et al. (Reference Ohlsson, Schlatter, Fischer and Henningson2010).

The computational grid used for the DNS had 58 720 spectral elements of order $p=6$ , giving 12.7 million grid points. The grid nearest the centreline contains a cylindrical region with a stretched Cartesian grid, similar to the grid used for a DNS of pipe flow in Nek5000 (El-Khoury et al. Reference El-Khoury, Schlatter, Noorani, Fischer, Breethouwer and Johansson2013). In the present grid, this region contains 128 elements over the axial cross-section, and extends from $r=0$ to the outer radius of the inner pipe at $r=0.14$ , where it attaches to an outer cylindrical grid. For $r>0.14$ , the grid is fully cylindrical with 32 elements over the circumference, which leads to a denser element distribution closer to the centreline where most of the interesting dynamics occur, and a coarser element distribution near the outer wall of the combustion chamber.

Figure 2.

Vorticity (helicity) from DNS at different axial cross-sections. The mesh is shown in red on top. This shows that the vorticity is continuous across the element boundaries, which is a sign of adequate resolution in spectral element method simulations of turbulent flows. (a) Typical contour upstream in the chamber ( $z=0.2$ ). (b) Typical contour downstream in the chamber ( $z=4$ ). (c) Contours at $z=4$ in a different colourscale (see colourbar), emphasizing regions of weak helicity.

This study focuses on the large-scale structures, especially the precessing vortex core, so detailed turbulence statistics such as wavenumber spectra (which would require storing a huge number of snapshots) have not been computed. To ensure that the resolution is sufficient for the task at hand, the following checks have been made on the nonlinear simulation results. First, we investigated the sensitivity of the precessing vortex core to the mesh resolution by decreasing the polynomial order inside each element. Going from $p=6$ to $p=5$ (42 % less degrees of freedom), the frequency from PSD signals of the precessing vortex core remained unaltered at $St=0.67$ . Only when going from $p=6$ to $p=4$ (70 % less degrees of freedom), did the frequency change slightly to $St=0.70$ . The frequency obtained with $p=6$ is also exactly the same as in experiments of Midgley et al. (Reference Midgley, Spencer and McGuirk2005) at much higher Reynolds number, confirming that the same physics is present. Furthermore, instantaneous velocity contours (showing the precessing vortex core) were confirmed to remain qualitatively the same with the coarser mesh. Finally, we investigated to what extent the turbulent motions are captured. In spectral element methods, the derivatives are not continuous across the spectral element boundaries, but become very close to continuous with increasing resolution. Therefore, in DNS using spectral element methods, a good indicator of adequate resolution of the turbulence is whether or not the vorticity or helicity fields look continuous across element boundaries (e.g. El-Khoury et al. Reference El-Khoury, Schlatter, Noorani, Fischer, Breethouwer and Johansson2013). Figure 2 shows typical contours of the helicity in the same mesh used in the DNS in this paper. No discontinuities are seen across the element boundaries. Furthermore, the vorticity shows the expected physical trends. Upstream in the chamber (figure 2 a), we observe fine-scale vorticity, particularly in the vortex core, surrounded by a thin ring of vorticity originating at the inlet wall ( $r=0.5$ ). Near the downstream wall of the chamber (figure 2 b), the vorticity has larger scales and is smoothly distributed along the whole radial extent, although it is still strongest near the centre.

The mean flows in the present study were computed by continuously time-averaging the velocity fields at every 10th time step over a time period of $150$ non-dimensional units. The mean flow was also averaged over the azimuthal direction, by interpolating the values at every grid point $(z_{n},r_{n},{\it\theta}_{n})$ into $(z_{n},r_{n},k2{\rm\pi}/32)$ , $k=1,\ldots ,32$ , and taking the average over all $k$ .

The POD modes were computed based on two different series of snapshots from DNS as follows. First, a series of 153 snapshots over a long time interval, $T=153$ , was saved and used to obtain the spatial shapes. A long time interval between the snapshots, ${\rm\Delta}t=0.5$ , was chosen to make them statistically independent. Second, a shorter series of frequently spaced snapshots, ${\rm\Delta}t=0.03$ apart, was used to obtain the mode frequencies. In both cases the procedure was as follows. First, the mean velocity field $\overline{\boldsymbol{U}}$ (obtained earlier) was subtracted from every snapshot. Subtracting the mean flow before performing the POD removes the mean flow mode, which otherwise would be the highest-energy mode. A similar zero-frequency mode will however be obtained at a lower energy, representing the difference between the mean flow averaged over all time steps, and the average over a finite number of snapshots. The snapshots with mean flow subtracted were then saved only on the part of the grid extending from $z=-1$ to $z=4$ in the axial direction, and with a lower polynomial order $p=4$ . This was done in order to save memory so that postprocessing in Matlab became possible, and in order to focus on the region where the coherent structures were visible. Next, the snapshots were uploaded into Matlab and a matrix formed with these columns:

(4.1) $$\begin{eqnarray}\unicode[STIX]{x1D653}\unicode[STIX]{x1D653}=[\boldsymbol{U}(t_{0})-\overline{\boldsymbol{U}},\boldsymbol{U}(t_{0}+{\it\delta}t)-\overline{\boldsymbol{U}},\ldots ,\boldsymbol{U}(t_{0}+T)-\overline{\boldsymbol{U}}].\end{eqnarray}$$

The POD modes were obtained from the singular value decomposition of the snapshot matrix: $\unicode[STIX]{x1D653}\unicode[STIX]{x1D653}=\boldsymbol{U}\unicode[STIX]{x1D64E}\boldsymbol{V}^{\text{T}}$ , where $\boldsymbol{U}$ contains the POD modes, the energies are obtained using $\unicode[STIX]{x1D64E}$ , and the time coefficients are $\boldsymbol{a}=\unicode[STIX]{x1D64E}\boldsymbol{V}^{\text{T}}$ .

4.2 Extraction of the eddy viscosity from POD

To compute the turbulent Reynolds stresses in DNS, we used the already computed POD modes from the long time series. First, the snapshot matrix $\unicode[STIX]{x1D653}\unicode[STIX]{x1D653}$ was reconstructed while setting to zero the time coefficients for the first four POD modes, which represent coherent structures, and the fifth POD mode, which represents the mean flow:

(4.2) $$\begin{eqnarray}\unicode[STIX]{x1D653}\unicode[STIX]{x1D653}_{new}=\mathop{\sum }_{k=6}^{N}\boldsymbol{U}(\boldsymbol{ : },k)\boldsymbol{a}(k,\boldsymbol{ : }).\end{eqnarray}$$

Now, every column of $\unicode[STIX]{x1D653}\unicode[STIX]{x1D653}_{new}$ is a snapshot of the original time series excluding the coherent structures, which we label a ‘reduced snapshot’. Second, the Reynolds stresses were computed and an average taken over all reduced snapshots. Third, the stresses were transformed to cylindrical coordinates and averaged over the azimuthal direction in the same way as the mean flow. Finally, they were interpolated back to Cartesian coordinates, and combined with mean flow stresses to form the product shown in (3.7). It was observed that the ratio between the Reynolds stresses and mean flow stresses can become ill-conditioned in regions where both of them are extremely small. Hence, a cutoff of $10^{-4}$ was employed on both stresses, and a cutoff of 100 was employed on the turbulent viscosity.

4.3 Linear global modes

We implemented the global cylindrical linear stability equations with an azimuthal wavenumber $m$ and a variable viscosity using the discretization and coding environment provided by FreeFem++ in combination with ARPACK, in the same way as Tammisola et al. (Reference Tammisola, Giannetti, Citro and Juniper2014) for a planar Cartesian geometry. The adjoint stability equations were also derived and implemented (continuous adjoint approach) in the case of molecular viscosity. For the variable viscosity case, the adjoint equations were not derived explicitly, but the adjoint modes were obtained using the conjugate transpose of the direct system matrix (discrete adjoint approach).

The spatial domain was discretized by a triangular finite elements mesh using a Delaunay–Voronoi algorithm, leading to a mesh with 213 620 triangles and 108 486 vertices. We employed the pair $\mathbb{P}_{2}{-}\mathbb{P}_{1}$ , consisting of piece-wise quadratic velocities and piece-wise linear pressure (Taylor–Hood elements), leading to $10^{6}$ degrees of freedom. The Nek5000 code was also used to compute modes without eddy viscosity in appendix A. The eigenmodes in Nek5000 were computed directly from the ansatz $\boldsymbol{u}(x,y,z,t)=\hat{\boldsymbol{u}}(x,y,z)\exp ({\it\sigma}t)$ , without setting an azimuthal wavenumber. This means that the obtained eigenspectrum from Nek5000 was a combination of all azimuthal wavenumbers. The two codes were cross-validated against each other, and the resolution independence of the eigenvalues tested, using molecular viscosity. The eigenvalues are documented in table 1 in appendix A.

Figure 3.

Mean flow from DNS at $Re=4800$ : (a) axial velocity, (b) radial velocity, (c) azimuthal velocity.

Figure 4.

Mean flow from DNS at $Re=4800$ , contours of the axial–radial velocity streamfunction.

Figure 5.

Instantaneous velocity from DNS at $Re=4800$ , axial velocity at a cross-section of the injector, at two different time instants (a and b). Light colours indicate high velocity positive in the streamwise direction, and dark colours negative velocity in the streamwise direction. Contours of (half of) the injector are also visible.

Figure 6.

(a) Illustration showing the location of the probe signals A, B and C in the $z$ – $r$ plane (eight other probes were also recorded). (b) PSD spectrum of probe A at $z=0.25$ , $r=0$ , (c) PSD spectrum of probe B at $z=0.05$ , $r=0.35$ ( ${\it\theta}={\rm\pi}/4$ ) (d) PSD spectrum of probe C at $z=1$ , $r=1.5$ ( ${\it\theta}={\rm\pi}$ ).

Figure 7.

POD mode 1 and mode 3 from the two leading mode pairs (mode 2 is the same as mode 1 with ${\rm\pi}/2$ phase difference, and mode 4 is the same as mode 3 with ${\rm\pi}/4$ phase difference): (a) mode 1, 3D contour of axial velocity; (b) the same for mode 3; (c) mode 1, axial velocity at the azimuthal cross-section at ${\it\theta}={\rm\pi}/2$ ; (d) the same for mode 3; (e) mode 1, the same as (c) but filtered by FFT in the azimuthal direction; (f) the same for mode 3.

Figure 8.

The leading Fourier component only ( $m=1$ ) for the first POD mode pair: mode 1 (a,c,e) and mode 2 (b,d, f). (a,b) Axial, (c,d) radial and (e,f) azimuthal velocity.

Figure 9.

The leading Fourier component only ( $m=2$ ) for the second POD mode pair: mode 3 (a,c,e) and mode 4 (b,d, f). (a,b) Axial, (c,d) radial and (e,f) azimuthal velocity.

Figure 10.

(a) The FFT coefficients of the chronos (time-domain) part of the first 11 POD modes, multiplied by their respective energies (see legend for lines/markers). The figure illustrates that POD mode pair 1–2 (blue online) is relatively monofrequent, and also is the by far most energetic coherent structure in the flow. A weaker coherent structure is seen in POD mode pair 3–4 (red online). (b) The same but shown in a logarithmic scale on both axes.

6 Turbulent viscosity from DNS

To approximate the effect of turbulent dissipation and its modelling on the linear global modes in the next section, we have extracted an approximate eddy-viscosity distribution from our DNS data using the simple Newtonian eddy model, as very recently used by Oberleithner et al. (Reference Oberleithner, Stöhr, Im, Arndt and Steinberg2015), and also suggested by Mettot et al. (Reference Mettot, Sipp and Bézard2014b ). The idea is based on a triple decomposition of the turbulent flow field:

(6.1) $$\begin{eqnarray}\boldsymbol{u}_{tot}=\boldsymbol{U}+\widetilde{\boldsymbol{u}}+\boldsymbol{u}^{\prime },\end{eqnarray}$$

where $\boldsymbol{U}$ is the mean flow, $\widetilde{\boldsymbol{u}}$ are the large-scale coherent structures, and $\boldsymbol{u}^{\prime }$ the small-scale turbulent fluctuations, which we assume can be modelled by an eddy viscosity (§ 3.1). To extract the latter from the DNS data, we first estimated that the large-scale coherent structures $\widetilde{\boldsymbol{u}}$ would be represented by the first four POD modes (the $m=1$ and $m=2$ modes analysed in § 5.1). We then reconstructed the flow field using all the other POD modes while setting the time coefficients for the first four to zero, to obtain $\boldsymbol{u}^{\prime }$ . Further, we extracted an isotropic turbulent viscosity as detailed in § 4.2. Contours of the turbulent viscosity (normalized by the molecular viscosity) are shown in figure 11. The turbulent viscosity is seen to be negligible in the upstream part of the nozzle and near the walls, while it is very high ( ${\it\mu}_{t}/{\it\mu}>10$ ) in most parts of the combustion chamber, where turbulence kinetic energy is high and mean flow stresses comparably low due to the expansion.

Figure 11.

Contours of ${\it\mu}_{t}/{\it\mu}$ , where ${\it\mu}_{t}$ is the turbulent viscosity extracted from the POD modes excluding the coherent structures (1–4). ${\it\mu}_{t}$ is used to form the effective Reynolds number ((3.9) and (3.7)). Observe that the contours are logarithmically spaced.

Figure 12.

The global linear eigenvalue spectrum around the DNS mean flow (figure 3) at $Re=4800$ : (a) modes with $m=1$ , (b) modes with $m=2$ .

Figure 13.

$m=1$ mode. (a,c,e) Global mode: (a) axial, (c) radial, (e) azimuthal velocity. (b,d, f) POD mode 1 from DNS: (b) axial, (d) radial and (f) azimuthal velocity.

Figure 14.

(a) Magnitude of the adjoint velocity of the global mode representing the precessing vortex core (the direct mode shown in figure 13 a,c,e). (b) Structural sensitivity of the same precessing vortex core mode.

7 Precessing vortex core as a global mode

We now seek to identify the precessing vortex core (PVC) as a global mode, with the aim of quantifying its structural sensitivity. Structural sensitivity can indicate where in the flow the PVC originates, and where the PVC may be influenced by changes in the flow and geometry.

The linear global modes were computed in FreeFem++ around the mean flow obtained from DNS. The whole eigenvalue spectra were first computed with different azimuthal wavenumbers on the coarser mesh with $5.5\times 10^{5}$ degrees of freedom. The computation was then repeated using a shift around the dominant mode, on a finer mesh with $1.23\times 10^{6}$ degrees of freedom, and the finer mesh was used to obtain results for eigenmode shapes and wavemakers. To include the effect of turbulent dissipation on the eigenmodes, the turbulent viscosity (figure 11) was used to generate a spatially varying Reynolds number $Re_{eff}(z,r,{\it\theta})$ . The effective Reynolds number was subsequently used in the stability computations. This is the approach in which the Reynolds stresses themselves are not linearized, named the ‘frozen eddy viscosity’ approach in § 1.

The global eigenvalue spectra computed this way are shown in figure 12 for azimuthal wavenumbers $m=1$ (a) and $m=2$ (b). To find oscillators, we turn our attention to the discrete global modes, separated from the continuous branch (which represents convective instabilities). Precisely one oscillator candidate is found: a $m=1$ mode with close to neutral growth rate. All other modes have very low growth rates. Modes at higher wavenumbers ( $m=3$ , $m=4$ ) have also been computed, and have an even lower growth rate.

To confirm that the selected global mode captures the correct physics, we compare the nearly neutral $m=1$ global mode with the most energetic POD mode pair, which was seen to be very close to monofrequent (figure 10). The frequency of the linear global mode is $St=0.74$ , which compares very well with the POD mode frequency: $St=0.7$ . The resulting direct global instability eigenmode with $m=1$ is shown in figure 13. The global mode is shown in (a,c,e), and the corresponding POD mode from DNS in (b,d,f) for comparison. The agreement with the POD mode shapes is excellent: the wavelength matches throughout the whole domain, and the shape and amplitude distribution also agree well.

With the goal of finding the wavemaker of the precessing vortex core more, its corresponding adjoint eigenmode has also been computed. The adjoint mode represents the optimal initial condition to excite the global mode (Chomaz Reference Chomaz2005). The magnitude of the adjoint velocity eigenmode is shown in figure 14(a). The adjoint mode is strongest inside the nozzle, along the whole extent of a shear/vorticity layer which starts at the inlet and includes the vortex breakdown location. This shows that any velocity perturbation that is on the streamline that impinges on the wavemaker region (discussed next) will have the strongest influence.

For oscillators such as the precessing vortex core, an optimal initial condition is not enough to alter the system dynamics. The eigenvalue needs to change, so the receptivity (adjoint mode) needs to overlap with a high amplitude of the direct mode. We will now overlap the direct and adjoint modes to obtain the structural sensitivity, given by the 2-norm of the structural sensitivity tensor. The structural sensitivity of the precessing vortex core is shown in figure 14(b). Knowing that the adjoint mode is strong upstream of the separation point, and the direct mode is strong downstream of the separation point, it is not surprising a posteriori that the wavemaker (structural sensitivity) resides near the separation point, which is the upstream point of the inner vortex breakdown bubble. Without prior knowledge, however, one might not have guessed that the structural sensitivity is so localized. The complicated mean flow with multiple recirculation zones (figure 4) might be expected to give rise to a range of different instability mechanisms. The sensitivity, however, points out a very specific region as the origin of the dominant instability: the vorticity layer and high-swirl region prior to the separation region near the nozzle exit.

Figure 15.

Profiles in the axial location of the maximum structural sensitivity inside the nozzle ( $z=-0.24$ ): (a) mean flow velocity profile, axial ( $U_{z}$ ) and swirl ( $U_{{\it\theta}}$ ) components. Structural sensitivity magnitude is shown by the greyscale on top. (b) Profile of the structural sensitivity magnitude in the same axial location.

8 Instability mechanism

We will now use the structural sensitivity to investigate which instability mechanisms are active in the precessing vortex core instability.

The axial location of the maximum structural sensitivity is inside the nozzle, at $z=-0.24$ . Mean flow velocity profiles in the axial location of the maximum sensitivity are shown in figure 15(a). Spiral vortex breakdown is an oscillatory motion which may occur in swirling flows that have a potential core and decaying swirl in the outer region (Leibovich Reference Leibovich1978), as in the mean swirl profile shown here. Spiral vortex breakdown occurs at swirl values slightly higher than that of the steady axisymmetric breakdown. Axisymmetric breakdown is the cause of flow separation inside a nozzle in swirl injectors (Leibovich Reference Leibovich1978; Syred Reference Syred2006). This flow (figure 3) separates in the nozzle and spiral vortex breakdown is possible.

The swirl number for spiral vortex breakdown is similar in different swirling flows: $S=0.915$ for a vortex breakdown bubble (Ruith et al. Reference Ruith, Chen, Meiburg and Maxworthy2003; Meliga et al. Reference Meliga, Gallaire and Chomaz2012a ) and $S=0.88$ for a swirling jet (Oberleithner et al. Reference Oberleithner, Sieber and Nayeri2011). In the position of the wavemaker in this flow, the swirl number based on the average velocities is $Sw=1.13$ . We suggest that the precessing vortex core is a spiral vortex breakdown instability, which takes place inside the nozzle in this injector flow mainly for two reasons: (i) the contraction leads to an increase of swirl through conservation of angular momentum until the swirl number reaches a critical value, and (ii) the favourable pressure gradient (which otherwise may suppress vortex breakdown, Leibovich Reference Leibovich1978) is less strong near the nozzle outlet where the flow starts to expand radially.

The structural sensitivity of the spiral vortex breakdown around an axisymmetric vortex breakdown bubble was studied by Qadri et al. (Reference Qadri, Mistry and Juniper2013). The structural sensitivity magnitude was high at the upstream end of the recirculation zone, with maximum amplitude at the centreline just upstream of the recirculation bubble. By considering different components of the structural sensitivity tensor, the instability mechanism was proposed to be due to the conservation of angular momentum upstream of the bubble, amplified by Kelvin–Helmholtz instability waves in the shear layer around the bubble.

In our flow, the structural sensitivity is also high just upstream of the separation point. However, the peak of the structural sensitivity in the vertical direction (light colours in figure 15) is in the shear layer. More tellingly, the peak of the structural sensitivity coincides with the inflection point in the shear layer. The whole profile of structural sensitivity magnitude as a function of vertical coordinate is shown in figure 15(b). The structural sensitivity of a parallel wake, shown in Qadri et al. (Reference Qadri, Mistry and Juniper2013), has a very similar appearance. This suggests that Kelvin–Helmholtz mechanism may be more influential for the injector flow than it is for the axisymmetric vortex breakdown bubble.

Figure 16.

Magnitude of components of the structural sensitivity tensor: (a) $u_{z}u_{z}^{+}$ , (b) $u_{z}u_{r}^{+}$ , (c) $u_{z}u_{{\it\theta}}^{+}$ , (d) $u_{r}u_{z}^{+}$ , (e) $u_{r}u_{r}^{+}$ , (f) $u_{r}u_{{\it\theta}}^{+}$ , (g) $u_{{\it\theta}}u_{z}^{+}$ , (h) $u_{{\it\theta}}u_{r}^{+}$ , (i) $u_{{\it\theta}}u_{{\it\theta}}^{+}$ . The colourscale goes from $0$ (dark) to $17.2$ (light).

In Qadri et al. (Reference Qadri, Mistry and Juniper2013), the relative importance of angular momentum conservation and Kelvin–Helmholtz instability was investigated based on the relative magnitudes of individual components of the structural sensitivity tensor $\unicode[STIX]{x1D61A}_{ij}=u_{i}u_{j}^{+}$ , where the index represents the axial, radial or azimuthal velocity component of the eigenmode $u$ or its adjoint $u^{+}$ . For example, the component $\unicode[STIX]{x1D61A}_{zr}$ represents a physical feedback mechanism where the radial momentum equation is perturbed by a force proportional to $u_{z}$ (for example, axial drag). The magnitude $\Vert \unicode[STIX]{x1D61A}_{zr}\Vert$ represents the maximal eigenvalue change caused by such a mechanism. The magnitudes of the nine components of $\unicode[STIX]{x1D61A}_{ij}$ are shown in figure 16(a–i). The $\unicode[STIX]{x1D61A}_{zz}$ component dominates, as was shown to be the case for the Kelvin–Helmholtz instability of a parallel wake in Qadri et al. (Reference Qadri, Mistry and Juniper2013). In contrast to the axisymmetric vortex breakdown, the feedback from angular momentum is weaker in comparison (components $\unicode[STIX]{x1D61A}_{r{\it\theta}}$ , $\unicode[STIX]{x1D61A}_{rr}$ , $\unicode[STIX]{x1D61A}_{{\it\theta}r}$ , $\unicode[STIX]{x1D61A}_{{\it\theta}{\it\theta}}$ ). This further confirms that the frequency of the global mode is selected at the shear layer inflection point by a pure Kelvin–Helmholtz mechanism. The shear layer inflection point in turn is caused by the axisymmetric vortex breakdown, and appears upstream of the breakdown (separation). This also appears to be the frequency selection of the final oscillation in the system where nonlinearities and turbulence are fully developed.

Figure 17.

Global mode derived by including eddy viscosity (a,c) compared to the same global mode derived without eddy viscosity (b,d): (a) $m=1$ axial velocity, with eddy viscosity, (b) $m=1$ , axial velocity, with molecular viscosity, (c) structural sensitivity, with eddy viscosity, (d) structural sensitivity, with molecular viscosity.

9 Effects of excluding the turbulent dissipation on the eigenmodes

In figure 17, the axial velocity of the global mode with eddy viscosity (a) is shown alongside the leading global mode computed around the mean flow, but with only molecular viscosity (b). The latter is the quasi-laminar approach used in Mettot et al. (Reference Mettot, Sipp and Bézard2014b ). In the mode shape with molecular viscosity, we can see the boundary of the central recirculation zone, showing that the mode shape and inclination is sensitive to the local shear. In contrast, the mode shape with eddy viscosity (figure 17 a) contains vertically aligned smooth structures extending inside and outside of the recirculation zone, exactly like the POD mode. We also see that the wavelength of the global mode with molecular viscosity shortens downstream, and the mode has still a visible amplitude at $z=4$ , while the wavelength of the mode with eddy viscosity does not shorten and the mode disappears around $z=3$ , in agreement with the POD mode.

There is little doubt that eddy viscosity improves the agreement of the global mode shapes with the POD, and the same seems to apply to the frequencies. The mode with eddy viscosity has $St=0.74$ , compared to $St=0.70\pm 0.06$ from DNS, and $St=0.87$ for the mode with molecular viscosity. Hence, eddy viscosity also notably improves the frequency agreement. The only drawback of the eddy viscosity is noted when looking at the mode growth rates: the stabilizing influence of turbulent dissipation is overestimated by making all eigenmodes stable. The $m=1$ mode is slightly stable with amplification rate ${\it\sigma}_{r}=-0.17$ . This highlights the issue known from previous studies that it may be difficult to obtain neutral growth rates in a mean flow analysis, and this might be a problem in cases where the growth rate is used as an indicator to identify the dominant eigenmodes. It should be mentioned that the mean flow stresses used to determine the eddy-viscosity distribution include the effects from both the turbulent fluctuations and the coherent structures. To be fully consistent, the base flow for a stability computation with eddy viscosity should include the effect of turbulent dissipation but not dissipation by coherent structures (termed base flow approach in Mettot et al. (Reference Mettot, Sipp and Bézard2014b )). A new base flow forced only by the turbulent Reynolds stresses could be computed, and this would lead to a somewhat smaller eddy viscosity, and possibly closer to neutral mode growth rates.

Finally, as the eddy viscosity influences the mode shapes, and the mode shapes are used to construct the structural sensitivity, the eddy viscosity also has some influence on the structural sensitivity (wavemaker). The wavemakers with and without eddy viscosity are compared in figure 17(c,d). By comparing the wavemaker with eddy viscosity (c) and that with molecular viscosity (d), we see that the eddy viscosity has two effects: it makes the wavemaker more dispersed, and lifts it up slightly from the wall. Hence, the wavemaker with eddy viscosity has the same amplitude over a region starting just upstream of the separation point and ending just beyond the injector lip, while the wavemaker without eddy viscosity is more focused in the immediate vicinity of the separation point. Although the amplitude of the structural sensitivity is strongly influenced by the viscosity model used, the location of its maximum (the wavemaker region) does not change. It is inside the nozzle in the upstream part of the central recirculation zone.

The structural sensitivity gives the influence of a local perturbation of the system matrix on the eigenvalue, whether the perturbation comes from a physical or a numerical origin. Hence, apart from estimating the physical origin of the instability, the structural sensitivity also has a numerical interpretation. We can observe that eddy viscosity reduces the maximum amplitude of the structural sensitivity of the $m=1$ mode by an order of magnitude, from approximately 20.5 to approximately 2.5, showing that the eigenmode with molecular viscosity is more sensitive to perturbations. From the numerical solution point of view, this means that the modes without eddy viscosity, and the underlying mean flow, need to be highly resolved near the nozzle, and in particular in the region near the separation point. This makes sense based on what is known about high-Reynolds-number flows. However, the modes with eddy viscosity have a lower effective Reynolds number, and hence a more even and weaker structural sensitivity. Therefore, the use of eddy viscosity makes the global mode problem better conditioned numerically.